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C. Grosche
F. Steiner
Handbook of Feynman Path Integrals
~ Springer
Dr. Christian Grosche Universit~it H a m b u r g II. Institut ftir Theoretische Physik Luruper Chaussee 149 D22761 H a m b u r g Email:
[email protected] Professor Dr. Frank Steiner Universit~it U  m Abteilung Theoretische Physik AlbertEinsteinAllee n D89o69 Ulm Email:
[email protected] Physics and Astronomy
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ISSN oo813869 ISBN 354o571353 SpringerVerlag
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CataloginginPublication Data applied for Die Deutsche Bibliothek  CIPEinheitsaufnahme Grosche, Christian: Handbook of Feynman path integrals / C. Grosche; F. Steiner.  Berlin; Heidelberg; New York; Barcelona; Budapest; Hong Kong; London; Milan; Paris; Singapore; Tokyo: Springer, 1998 (Springer tracts in modern physics; Vol. 145) ISBN 3540571353 This work is subject to copyright. All rights are reserved, whether the whole or part of the material is concerned, specifically the rights of translation, reprinting, reuse of illustrations, recitation, broadcasting, reproduction on microfilm or in any other way, and storage in data banks. Duplication of this publication or parts thereof is permitted only under the provisions of the German Copyright Law of September 9,1965, in its current version, and permission for use must always be obtained from SpringerVerlag. Violations are liable for prosecution under the German Copyright Law. © SpringerVerlag Berlin Heidelberg 1998 Printed in Germany The use of general descriptive names, registered names, trademarks, etc. in this publication does not imply, even in the absence of a specific statement, that such names are exempt from the relevant protective laws and regulations and therefore free for general use. Typesetting: Cameraready copy by the authors using a Springer TEX macro package Cover design: design 6" production GmbH, Heidelberg SPIN: lo717u2 56/3o12  5 4 3 2 1  Printed on acidfree paper
To the memory of Isabel Steiner
Preface
Our Handbook of Feynman Path Integrals appears just fifty years after Richard Feynman published his pioneering paper in 1948 entitled "SpaceTime Approach to NonRelativistic Quantum Mechanics". As it is the case with many books, its origin goes back to a course first given by one of us (F.S.) on Feynman path integrals at the University of Hamburg during the summer semester of 1983. The other author was one of the students attending these lectures and who eventually decided to work on this subject for his diploma thesis. This was the starting point of our collaboration during the 1980s. At that time our main common interest was in the question of how to solve nonGaussian path integrals (like the one for the hydrogen atom) and, more generally, path integrals in arbitrary curvilinear coordinates. It was in 1983, too, that one of us (F.S.) began to collect papers and preprints on path integrals, and to set up a comprehensive list of references on this subject. Eventually a systematic literature search was carried out (by C.G.). While we were working in various fields, above all in quantum chromodynamics, string theory, and quantum chaos, we conceived the idea of writing a Handbook on Feynman path integrals which would, on the one hand, serve the reader as a thorough introduction to the theory of path integrals, but would, on the other hand, also establish for the first time a comprehensive table of Feynman path integrals together with an extensive list of references. The whole enterprise was, however, delayed by various circumstances for several years. Here we put forward our Handbook to the gentle reader! The book follows the general idea as originally conceived. Chapters 15 have the character of a textbook and give a selfcontained, and uptodate introduction to the theory of path integrals for those readers who have not yet studied path integrals, but have a good knowledge of the fundamentals of quantum mechanics as covered by standard courses in theoretical physics. Chapter 6 makes up the largest part of this Handbook and contains a rather complete table of path integrals in nonrelativistic quantum mechanics, including supersymmetric quantum mechanics, and statistical mechanics. To each path integral listed in the table we attach a comprehensive list of references which altogether make up almost 1000 references. The Introduction in Chap. 1 is mainly of a historical nature and gives the reader some insight into the remarkable development of Feynman's path integral approach. Since some of the historical facts are not so well known we thought it would be worthwhile to present them in Chap. 1.
VIII
Preface
Large parts of the material presented in Chaps. 15 have been used and tried out first by one of us (F.S.) in various courses given over the last 15 years at the Universities of Hamburg and Ulm, and at the University of Lausanne and the ETH Lausanne, respectively, in 1985 and 1995, as part of the Troisi~me Cycle de la Physique en Suisse Romande. We are grateful to all students and colleagues who have attended these lectures and who have contributed by their questions and remarks to the clarification and improvement of our presentation. We are indebted for help and criticism to many friends and colleagues, including Sergio Albeverio, Jens Bolte, Philippe Choquard, Ludwig D~browski, Gianfausto Dell'Antonio, Josef Devreese,
[email protected] DeWittMorette, Ismael Duru, Klaus Fredenhagen, Martin Gutzwiller, Urs Hugentobler, Akira Inomata, Chris Isham, Georg Junker, John Klauder, Hagen Kleinert, Hajo Leschke, Gerhard Mack, Dieter Mayer, Peter Minkowski, Holger Ninnemann, David Olive, George Papadopoulos, Axel Pelster, George Pogosyan, Cesare Reina, Martin Reuter, Oliver Rudolph, Virulh Sayakanit, Larry Schulman, Alexei Sissakian, Oleg Smolyanov, Wichit Sritrakool, Ulrich Weiss, Frederik Wiegel, Pavel Winternitz, Kurt Bernardo Wolf, and Arne Wunderlin. Important software advise was provided by Michael Behrens, Otto Hell, Phillip Kent, Dennis Moore, Jan Hendrik Peters, Peter Schilling, Thomas SippelDau, and Katherine Wipf. We are also grateful to our secretaries Ingrid Gruhler, Doris Laudahn, B~rbel Lossa, Graziella Negadi, Alexandra Poretti, and Galina Sandukovskaya. Furthermore we thank SpringerVerlag, in particular Urda and Wolf BeiglbSck, Hans KSlsch, and Jaequeline Lenz for their editorial guidance. Financial support by the Deutsche Forschungsgemeinschaft DFG is gratefully acknowledged. Last but not least C.G. is deeply indebted to Gertrude Huber and Diana Paris for their love, understanding and support at a critical time. It so happens that Feynman would have celebrated his 80th birthday on May 11 this year, and it seems therefore that the publication of our handbook is quite well timed. Hamburg and Ulm, May 1998
Christian Grosche Frank Steiner
Table
of Contents
1 Introduction
2 General
........................................................
Theory
...................................................
1 23
2.1 The Feynman Kernel and the Green Function . . . . . . . . . . . . . . . . . . . . .
23
2.2 The Path Integral in Cartesian Coordinates . . . . . . . . . . . . . . . . . . . . . . .
31
2.3 Gaussian Path Integrals and Zeta Function Regularization . . . . . . . . 37 2.4 Evaluation of Path Integrals by Fourier Series . . . . . . . . . . . . . . . . . . . . .
39
2.5 Path Integration Over Coherent States . . . . . . . . . . . . . . . . . . . . . . . . . . . .
44
2.6 Fermionic Path Integrals . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . .
55
2.7 The Path Integral in Spherical Coordinates . . . . . . . . . . . . . . . . . . . . . . .
63
2.8 The Path Integral in General Coordinates . . . . . . . . . . . . . . . . . . . . . . . .
67
2.9 Transformation Techniques . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . .
76
2.10 Exact Path Integral Treatment of the Hydrogen Atom . . . . . . . . . . .
87
2.11 The Path Integral in Parabolic Coordinates . . . . . . . . . . . . . . . . . . . . . .
88
3 Basic Path Integrals ..............................................
93
3.1 The Free Particle . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . .
94
3.2 The Quadratic Lagrangian . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . .
95
3.3 The Radial Harmonic Oscillator . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . .
99
3.4 Path Integration Over Group Manifolds . . . . . . . . . . . . . . . . . . . . . . . . . .
102
4 Perturbation
Theory ............................................
4.1 Path Integration and Perturbation Theory . . . . . . . . . . . . . . . . . . . . . . .
123 124
4.2 Summation of the Perturbation Series for (f and ~Potentials . . . . 127 4.3 Partition Functions and Effective Potentials . . . . . . . . . . . . . . . . . . . . .
132
4.4 Semiclassical Expansion About the Harmonic Approximation . . . . 134
X
Table of Contents
5 Semiclassical Theory
............................................
141
5.1 Semiclassical T h e o r y and Q u a n t u m Chaos . . . . . . . . . . . . . . . . . . . . . . .
141
5.2 Semiclassical Expansion of the F e y n m a n P a t h Integral . . . . . . . . . . .
143
5.3 Semiclassical Expansion of the Green Function . . . . . . . . . . . . . . . . . . .
149
5.4 The Gutzwiller Trace F o r m u l a . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . .
151
6 Table of Path Integrals .........................................
155
6.1 General Formulae . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . .
162
6.2 The General Q u a d r a t i c Lagrangian . . . . . . . . . . . . . . . . . . . . . . . . . . . . . .
173
6.3 Discontinuous Potentials . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . .
217
6.4 The R a d i a l Harmonic Oscillator . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . .
225
6.5 The PSschlTeller Potential . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . .
240
6.6 The Modified PSschlTeller Potential . . . . . . . . . . . . . . . . . . . . . . . . . . . .
244
6.7 Motion on G r o u p Spaces and Homogeneous Spaces . . . . . . . . . . . . . .
258
6.8 C o u l o m b Potentials . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . .
277
6.9 Magnetic Monopole and Anyon Systems . . . . . . . . . . . . . . . . . . . . . . . . .
295
6.10 Motion in Hyperbolic G e o m e t r y . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . .
304
6.11 Explicit T i m e  D e p e n d e n t Problems . . . . . . . . . . . . . . . . . . . . . . . . . . . . .
322
6.12 Point Interactions . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . .
327
6.13 B o u n d a r y Value Problems . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . .
347
6.14 Coherent States . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . .
353
6.15 Fermions . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . .
361
6.16 S u p e r s y m m e t r i c Q u a n t u m Mechanics . . . . . . . . . . . . . . . . . . . . . . . . . . .
363
References ..........................................................
367
List of Symbols
....................................................
425
Subject Index
......................................................
429
Author
......................................................
439
Index
1 Introduction
The conventional formulation of q u a n t u m mechanics in terms of operators in Hilbert space is a Hamiltonian approach) It was invented and developed by Bohr, Born, Dirac, Heisenberg, Jordan, Pauli, SchrSdinger, and others in the years 192526. The basic quantity in q u a n t u m mechanics is a certain complex function kV called a probability amplitude or wave function associated with every q u a n t u m mechanical (pure) state. In the simplest case of a single particle the wave function kV(x,t) is the total amplitude for the particle to arrive at a particular point (x, t) in space and time from the past in some (perhaps unspecified) situation. The probability (density) of finding the particle at the point x and at the time t is [k~(x,t)[ 2. In the usual approach to q u a n t u m mechanics the wave function kv is calculated by solving a differential equation, which for nonrelativistic systems, i.e., for particles of low velocity, is the SchrSdinger equation
i hOgr'g7 ' t) _ H_x,P(x, t) .
(1.1.1)
Here I:Ix is a differential operator called the Hamiltonian or SchrSdinger operator, which is derived from the classical Hamiltonian H ( p , x) of the associated classical system. 2 The SchrSdinger equation (1.1.1) is a kind of wave equation, a and this explains why the probability amplitude ~P(x, t) is commonly called the (SchrSdinger) wave function. Obviously, the SchrSdinger equation (1.1.1) is a deterministic equation, since knowledge of ~P at t = t' 1 In the following discussion we shall not distinguish between Heisenberg's matrix mechanics discovered in June 1925 and SchrSdinger's undulatory mechanics discovered during the winter 192526, since the two, apparently dissimilar approaches, were proved to be mathematically equivalent by SchrSdinger, and independently by Dirac, already in 1926. 2 For details, the reader is referred to Sect. 2. 3 The crucial point that the factor i = x/rZ'i" in (1.1.1) is unavoidable, took Schr6dinger several months to finally accept. See the discussions in [867, 938] on this point. It was recognized by Ehrenfest in 1927 that an analytic continuation in time to "Euchdean time", t ~  i t, transforms the SchrSdinger equation into the heat or diffusion equation [295], see also the remarks on the FeynmanKac formula on p.18 and in Sect. 2.2.
2
Introduction
implies its knowledge at all subsequent times t H > t ~. However, the interpretation of I~Pl~ as the probability of an event" is an indeterministic interpretation. 4 Introducing the Green function K(x",x~;t",tO of the timedependent Schrbdinger equation (1.1.1), the quantum mechanical time evolution of the wave function ~(x, t) is explicitly given by the integral relation # ( x " , t " )  J dx' g(x",x';t",t')#(x',t')
,
(1.1.2)
which determines the probability amplitude at a final point x" at time t" in terms of the probability amplitude #'(x p, t ~) at an initial point x ~ at time tq Equation (1.1.2) shows that the Green function K plays the r61e of an integral kernel. In fact, K is identical to the kernel of the quantum mechanical timeevolution operator (T = t"  t ~ > 0) i
H
see also (2.1.19). Since the integral relation (1.1.2) is completely equivalent to the Schrbdinger equation (1.1.1), it offers the possibility of considering (1.1.2) as the basic timeevolution equation in quantum mechanics and thus as an alternative to the operator Schrbdinger equation. This is exactly Feynman's approach in his path integral formulation of quantum mechanics [326]. In this approach the integral kernel K is the primary object, and that is the reason why the timedependent Green function K is in this context commonly called the Feynman kernel. 5 "A quantum mechanical system is described equally well by specifying the function K, or by specifying the Hamiltonian H x from which it results. For some purposes the specification in terms of K is easier to use and visualize." [328]. It is clear from (1.1.2) and (1.1.3) that the Feynman kernel K (x ~, x~; t", t ~) has the meaning of a transitionprobability amplitude to get from the point (x',t') to the point (x",t"), or in Feynman's words: "A probability amplitude is associated with an entire motion of a particle as a function of time, rather than simply with a position of the particle at a particular time" [326]. It is a remarkable fact that by taking the Feynman kernel K as the primary object, one is led to a novel formulation of quantum mechanics (though mathematically equivalent to the more usual Hamiltonian approach) which turns out to be a Lagrangian formulation. Furthermore, instead of operators in Hilbert space and differential equations one has to deal with Feynman path integrals called functional integrals in mathematics. Although the path integral formulation to quantum mechanics, as a complete theory, is solely a For a very lucid exposition of the fundamental concepts of probability and probability amplitudes in quantum mechanics, see Feynman's original paper [326], the FeynmanHibbs book [340], and [669]. 5 This is the terminology which we shall adopt throughout this book. Note, however, that the Feynman kernel K is also called the propagator by some authors.
1 Introduction
3
the work of Feynman [325,326], the important discovery of the close analogy between the Feynman kernel K ( x " , xl; t J+ dr, t ~) associated with an infinitesimal displacement from time t ~ to time t~+ dt and the phase factor exp[~/2 dt] with s being the classical Lagrangian is due to Dirac [254]. By 1927 Dirac had worked out his transformation theory [252] by studying socalled quantum mechanical transformation functions which depend on pairs of conjugate variables which, at the classical level, are connected via canonical transformations. 6 However, the canonical transformations considered by Dirac in this paper come from a generating function which is of type 2 (following the traditional labeling of canonical transformations), i.e., it connects an initial m o m e n t u m P with a final position q. Thus the associated transformation function called (alP) by Dirac [252] cannot be identified with the Feynman kernel K which reads in Dirac's notation of 1932 (qt"lqt'). (In our notation we have K(x",x';t",t') = ( x " , t " l x ' , t ' ) with Ix, t) := e x p ( + ~ t _Hx)[x), see (1.1.3).) 7 The time transformation function (at" Iqt') appears for the first time in Dirac's paper [254] entitled "The Lagrangian in Quantum Mechanics" received by the Physikalische Zeitschrift der Sowjetunion on November 19, 1932. Let us quote from the introduction of this paper [254, pp. 64, 65]: "Quantum mechanics was built up on a foundation of analogy with the Hamiltonian theory of classical mechanics. This is because the classical notion of canonical coordinates and momenta was found to be one with a very simple quantum analogue, as a result of which the whole of the classical Hamiltonian theory, which is just a structure built up on this notion, could be taken over in all its details into quantum mechanics.  Now there is an alternative formulation for classical dynamics, provided by the Lagrangian. This requires one to work in terms of coordinates and velocities instead of coordinates and momenta. The two formulations are, of course, closely related, but there are reasons for believing that the Lagrangian one is the more fundamental . . . . For these reasons it would seem desirable to take up the question of what corresponds in the quantum theory to the Lagrangian method of the classical theory. A little consideration shows, however, that one cannot expect to be able to take over the classical Lagrangian equations in any very direct way. These equations involve partial derivatives of the Lagrangian with respect to the coordinates and velocities and no meaning can be given to such derivatives in quantum mechanics. The only differentiation process that can be carried out with respect to the dynamical variables of quantum mechanics is that of forming Poisson brackets and this process leads to the Hamiltonian theory . . . . We must therefore seek our quantum Lagrangian theory in an indirect way. We must try to take over the ideas of 6 The earlier work of Jordan [546] concerns timeindependent canonical transformations which are not relevant for our discussion. 7 The transformation function (qlP) is the subject of Van Vleck's famous paper [906]. For a discussion of this paper, its relation to Dirac's papers [252, 254] and the story of Van Vleck's determinant, see [193].
4
Introduction
the classical Lagrangian theory, not the equations of the classical Lagrangian theory." Diraz considers [254] two sets of conjugate variables (p, q) and (P, Q) but supposes now that, at the classical level, the independent variables of the generating function are q and Q. Let R be the corresponding generating function, s Then the corresponding dependent variables are given by
OR P=~q '
OR POQ '
(1.1.4)
where R = R(q, Q; t). Notice that this generating function of canonical transformations is of type 1. In the section entitled "The Lagrangian and the Action Principle" Dirac [1.c., p.67] continues: "The equations of motion of the classical theory cause the dynamical variables to vary in such a way that their values qt,Pt at any time t are connected with their values qT, PT at any other time T by a contact transformation, which may be put into the form (1.1.4) with q,p = qt,Pt; Q, P = qT, PT and R equal to the time integral of the Lagrangian over the range T to t. In the quantum theory the qt,Pt will still be connected with the qT,PT by a contact transformation and there will be a transformation function (qt[qw) connecting the two representations in which the qt and the qT are diagonal respectively. The work of the preceding section now shows that
(qt[qT)
corresponds to
exp
s
,
(1.1.5)
where 1: is the Lagrangian. If we take T to differ only infinitesimally from t, we get the result
(qt+dtlqt)
corresponds to
exP ( h ~ d t )
.
(1.1.6)
The transformation functions in (1.1.5) and (1.1.6) are very fundamental things in the quantum theory and it is satisfactory to find that they have their classical analogues, expressible simply in terms of the Lagrangian." The above citations show quite clearly that Dirac had carried the programme of formulating a Lagrangian approach to quantum mechanics quite far. W h a t is strange, however, is that the "very fundamental things" appear in (1.1.5) and (1.1.6) not in the form of equations, but rather Dirac uses the vague formulation corresponds to. 9 We may thus ask with Schwinger [842]: "Why, then, did Dirac not make a more precise, if less general, statement? 8 Dirac uses S instead of R. 9 Already in the second edition of his famous book The Principles of Quantum Mechanics, which appeared in 1935 [256], Dirac included these results in Sect. 33, but neither in this nor in later editions did he quantify the correspondence expressed in (1.1.5) and (1.1.6). See also Dirac's paper from 1945 [255].
1 Introduction
5
Because he was interested only in a general question: What, in quantum mechanics, corresponds to the classical principle of stationary action?" In order to answer this general question, Dirac considered the basic composition law for (qtlqT) in the form (qt]qT) = ] ( q t [ q m ) d q m ( q r n [ q m  1 ) d q m  i . . . ( q ~ [ q l ) d q l ( q l [ q T )
,
(1.1.7)
where the time interval T + t has been divided up "into a large number of small sections T  + t l , t l + t ~ , . . . , tin1 + t,,,,t,~ + t by the introduction of a sequence of intermediate times tl, t2, 9 9 tin', and "where qk denotes q at the intermediate time tk, (k  1 , 2 , . . . , m ) . " He then compared the composition law (1.1.7) with the product A(tT) = A(ttm)A(tmtm1)...A(t2tl)A(tlT)
,
(1.1.8)
where
f's \_ = A ( t T )
exp ( ! \hJr
"so that A ( t T ) is the classical analogue of (qt[qT)." "Equation (1.1.7) at first sight does not seem to correspond to equation (1.1.8), since on the righthand side of (1.1.7) we must integrate after doing the multiplication while on the righthand side of (1.1.8) there is no integration. "Let us examine the discrepancy by seeing what becomes of (1.1.7) when we regard t as extremely small. From the results (1.1.5) and (1.1.6) we see that the integrand in (1.1.7) must be of the form e iF/h where F is a function of qT, ql, q2, 9 9 qm, qt which remains finite as h tends to zero. Let us now picture one of the intermediate q's, say qk, as varying continuously while the others are fixed. Owing to the smallness of h, we shall then in general have F / l i varying extremely rapidly. This means that e iF/h will vary periodically with a very high frequency about the value zero, as a result of which its integral will be practically zero. The only important part in the domain of integration of qk is thus that for which a comparatively large variation in qk produces only a very small variation in F. This part is the neighbourhood of a point for which F is stationary with respect to small variations in qk. "We can apply this argument to each of the variables of integration in the righthand side of (1.1.7) and obtain the result that the only important part in the domain of integration is that for which F is stationary for small variations in all the intermediate q's. But, by applying (1.1.5) to each of the small time sections, we see that F has for its classical analogue s m
+
s ml
s
+
s
=
s
,
(1.1.9)
dtx
which is just the action function which classical mechanics requires to be stationary for small variations in all the intermediate q's. This shows the way
6
Introduction
in which (1.1.7) goes over into classical results when h becomes extremely small." Thus Dirac showed by considering the semiclassical limit h + 0 that the multipleintegral construction (1.1.7) of the time transformation function contains the quantum analogue of the classical action principle, a fundamental result, indeed. "Why, in the decade that followed, didn't someone pick up the computational possibilities offered by this integral approach to the time transformation function? To answer this question bluntly, perhaps no one needed it until Feynman came along." [842]. Feynman 1~ was working as a research assistant at Princeton during 194041. In the course of his graduate studies he discovered together with Wheeler an action principle using half advanced and half retarded potentials [920]. The problem was the infinite selfenergy of the electron, and it turned out that the new action "principle could deal successfully with the infinity arising in the application of classical electrodynamics.  The problem then became one of applying this action principle to quantum mechanics in such a way that classical mechanics could arise naturally as a special case of quantum mechanics when h was allowed to go to zero.  Feynman searched for any ideas which might have been previously worked out in connecting quantummechanical behaviour with such classical ideas as the Lagrangian or, in particular, Hamilton's principal function R, the indefinite integral of the Lagrangian." [340]. At a Princeton beer party Feynman learned from Herbert Jehle, a former student of SchrSdinger in Berlin, who had newly arrived from Europe, of Dirac's paper [254]. The natural question that then arose was what Dirac had meant by the phrase "corresponds to", see (1.1.5) and (1.1.6). Feynman found that Dirac's statement actually means "proportional to" such that (1.1.6) is to within a constant factor an equality. Based on this result and the composition law (1.1.7) in the limit m 4 oo, Feynman interpreted the multipleintegral construction (1.1.7) as an "integral over all paths" and wrote this down for the first time in his Ph D thesis [325] presented to the Faculty of Princeton University on May 4, 1942. During the war Feynman worked at Los Alamos, and after the war his primary direction of work was towards quantum electrodynamics. So it happened that a complete theory of the path integral approach to quantum mechanics was worked out only in 1947. Feynman submitted his paper to The Physical Review, but the editors rejected it! Thus he rewrote it and sent it to Reviews of Modern Physics, where it finally appeared in spring 1948 under the title "SpaceTime Approach to NonRelativistic Quantum Mechanics" [326]. l~ the work and life of Richard P. Feynman, the reader may consult the following sources: Feynman's Nobel lecture [336]; Feynman's two autobiographies [338, 339]; the excellent biography by Gleick [396], including "a Feynman bibliography"; the special issue "Richard Feynman" in Physics Today 42 (February 1989); Dyson's autobiography [287], and Schweber's book [836].
1 Introduction
7
Feynman's paper [326] is one of the most beautiful and most influential papers in physics written during the last fifty years. While at first sight the path integral formulation appears to be "merely a reformulation of quantum mechanics, equivalent to the usual formulation" [377], there are now some indications that "the path integral formulation of quantum mechanics may be more fundamental than the conventional one, in that there is a crucial domain where it may apply and the conventional formulation may fail. That is the domain of quantum cosmology." [377] Let us briefly sketch how Feynman arrived [326] at his path integral) 1 First he considered the limit m + oo of the composition law (1.1.7), which is equivalent to the limit e + 0, if the intermediate times t k are, for simplicity, chosen to be equidistant, i.e., tk = t ' + kr (k = 0, 1 , . . . , N  1), with N := m + 1, t' := T = to, t" := t = tg, and c = (t"  t')/N. Then (1.1.7) becomes
(qt"lqt') = l i m / d q l . . . d q N  l ( q t " l q g  1 ) ( q g  l l q g  2 ) . . . e....+O J
(q2]ql)(qllqt')
(1.1.10) assuming that the limit exists, of course. Converting from Dirac's notation to the notation introduced in (1.1.2) and (1.1.3),one obtains
N1 K ( x " , x ' ; t " , t ' ) = ~+01imI T
k=l
N1
II
K(Xj+l,Xj;tj +e, tj)
.
(1.1.11)
j=0
This multipleintegral representation of the Feynman kernel is built up by the shorttime kernels K(xj+l, xj;tj + e,tj), (j = O, 1,..., N  1), for which Feynman [326] writes, in the limit e + 0,
thus replacing Dirac's vague correspondence (1.1.6) by a precise statement involving some normalization factor A = A(e) for each instant of time, suitably adjusted. (Remember that the Lagrangian depends on the trajectory x(t) and the velocity 5:(t), i.e., /: = s Inserting (1.1.12) into (1.1.11), Feynman obtained
K ( x " , x ' ; t " , t ' ) = ~+01imANH
/d~kexp
k=l
~ ~+1, x~+t ~J '=
(1.1.13) Since the sum in the exponent becomes in the limit r + 0 just the action R,
N1 es ( xj+l, xj+l  xj ) = ~tff' s
lim ~ e~. O
j=O
=: R[z(t)]
f
11A detailed account of the general theory will be given in Chap. 2.
(1.1.14)
8
Introduction
Feynman path integral ~(t")=~" K(x",x';t",t') = / 7Px(t)exP(hR[X(t)] ) . (1.1.15) x(t,)=x, x(t")=x" symbol f 79x(t)is defined 1~ by (1.1.13) and represents (some ~(t,)=~,
one arrives at the
Here the
kind of) an integration over the space of functions z(t), i.e., all possible paths, connecting the points (x',t') and ( x " , t " ) . For a particle of mass m moving in a onedimensional potential V(x), i.e., s = ~ x 2 _ V(x), Feynman derived for the normalization factor A A =
(1.1.16)
Many years later, in his Nobel lecture, Feynman described his discovery as follows [336]: "In that way I found myself thinking of a large number of integrals, one after the other in sequence. In the integrand was the product of the exponentials, which, of course, was the exponential of the sum of terms like e/;. Now L; is the Lagrangian and e is like the time interval dt, so that if you took a sum of such terms, t h a t ' s exactly like an integral. T h a t ' s like Riemann's formula for the integral f~.dt; you just take the value at each point and add t h e m together. We are to take the limit as e + 0, of course. Therefore, the connection between the wave function of one instant and the wave function of another instant a finite time later could be obtained by an infinite number of integrals (because e goes to zero, of course) of exponential (i R/h), where R is the action expression (1.1.14). At least, I had succeeded in representing q u a n t u m mechanics directly in terms of the action R . . . . This led later on to the idea of the amplitude of the path; that for each possible way that the particle can go from one point to another in spacetime, there's an amplitude. T h a t amplitude is e to the i / h times the action for the path. Amplitudes from various paths superpose by addition. This then is another, a third way, of describing q u a n t u m mechanics, which looks quite different than that of SchrSdinger or Heisenberg, but is equivalent to them." One of the first physicists who understood Feynman's "intuitive method" la was Dyson. He gave the following description of Feynman in those days [287]: "Dick was also a profoundly original scientist. He refused to take anybody's 12The identifying notation Dx(t) was not yet used in [326]. It was introduced by Feynman in [330]. 13"As a result", Feynman said [336], "the work was criticized, I don't know whether favorably or unfavorably, and the 'method' was called the 'intuitive method'. For those who do not realize it, however, I should like to emphasize that there is a lot of work involved in using this 'intuitive method' successfully . . . . Nevertheless, a very great deal more truth can become known than can be proven."
1 Introduction
9
word for anything. This meant that he was forced to rediscover or reinvent for himself almost the whole of physics. It took him five years of concentrated work to reinvent quantum mechanics. He said that he couldn't understand the official version of quantum mechanics that was taught in the textbooks, and so he had to begin afresh from the beginning. This was a heroic enterprise. He worked harder during those years than anybody else I ever knew. At the end he had his version of quantum mechanics that he could understand. He then went on to calculate with his version of quantum mechanics how an electron should behave. He was able to reproduce the result that Hans [Bethe] had calculated using orthodox theories a little earlier. But Dick could go much further. He calculated with his own theory fine details of the electron's behaviour that Hans's method could not touch. Dick could calculate these things far more accurately, and far more easily, than anybody else could. The calculation that I did for Hans, using the orthodox theory, took me several months of work and several hundred sheets of paper. Dick could get the same answer, calculating on a blackboard, in half an hour." In his last remarks on an electron's behaviour Dyson refers to Feynman's famous work in quantum electrodynamics, culminating in the Feynman rules [328330] and Feynman diagrams [329], which Feynman first derived 14 using the path integral method, and which nowadays can be found in every textbook on quantum field theory and elementary particle physics. Feynman describes this work as follows [336]: "The rest of my work was simply to improve the techniques then available for calculations, making diagrams to help analyze perturbation theory quicker. Most of this was first worked out by guessing  you see . . . . I included diagrams for the various terms of the perturbation series, improved notations to be used, worked out easy ways to evaluate integrals which occurred in these problems, and so on, and made a kind of handbook on how to do quantum electrodynamics . . . . At this stage, I was urged to publish this because everybody said it looks like an easy way to make calculations and wanted to know how to do it. I had to publish it missing two things; one was proof of every statement in a mathematically conventional sense. Often, even in a physicist's sense, I did not have a demonstration of how to get all of these rules and equations from conventional electrodynamics." The path integral (1.1.15) is the fundamental quantummechanical rule in Feynman's third way of describing quantum mechanics. The rule tells us [340] "how much each trajectory contributes to the total amplitude to go from (z', t') to (z", t"). It is not that just the particular path of extreme action contributes; rather, it is that all the paths contribute. They contribute equal amounts to the total amplitude, but contribute at different phases. The phase of the contribution from a given path is the action R for that path in units of the quantum of action h. That is, to summarize: The probability P(z",x';t",t ~) to go from a point z ~ at the time t' to the point x" at t" is the absolute square P(z", z'; t", t') = IK(z", z'; t", t')l 2 of an amplitude 14See Feynman's remark in [330].
10
Introduction
If(x", z'; t", t') to go from (x', t') to (x", t"). This amplitude is the sum of contributions O[x(t)] from each path K ( x " , x'; t", t') =
~
O[x(t)] .
(1.1.17)
over all paths from x ~ to x"
The contribution of a path has a phase proportional to the action R: O[x(t)] = const e (i/h)R[~(t)]
(1.1.18)
The action is that for the corresponding classical system, see (1.1.14)." The only purpose of rewriting the path integral (1.1.15) in the (even more symbolic) sum form (1.1.17) is to illustrate its interpretation as a sum over all paths or sum over all histories. Equation (1.1.17) makes it particularly clear that the total amplitude K depends on the whole spacetime history, i.e., all paths, and that it is obtained by a superposition of the amplitudes Cb[x(t)] from all paths x(t) which connect the spacetime points (x', t') and ( X " , t '#) .15
In this book we will mainly use the lattice definition (1.1.13) of the path integral (1.1.15). Almost all path integral solutions presented here have been obtained by using this definition (and its generalizations to many degrees of freedom, curvilinear coordinates, etc.), i.e., have been worked out using the subdivision and limiting processes involved in (1.1.13). Feynman was fully aware of the mathematical problems associated with the integration in functional spaces. Already in his 1948 paper he wrote in a footnote [326]: "There are very interesting mathematical problems involved in the attempts to avoid the subdivision and limiting processes. Some sort of complex measure is being associated with the space of functions x(t). Finite results can be obtained under unexpected circumstances because the measure is not positive everywhere, but the contributions from most of the paths largely cancel out. These curious mathematical problems are sidestepped by the subdivision process. However, one feels as Cavalieri must have felt calculating the volume of a pyramid before the invention of calculus." In writing the path integral (1.1.13) in the "less restrictive notation" (1.1.15), Feynman expressed his strong belief that "the concept of the sum over all paths, like the concept of an ordinary integral, is independent of a special definition and valid in spite of the failure of such definitions" [340]. At this point we will not discuss the mathematical aspects of path integrals, but rather we will focus our attention on the question of whether the new formulation of quantum mechanics had a favorable reception. It is not too surprising to learn that Feynman's ideas were not appreciated in the beginning among the physicists of the older generation who had laid lSActually, it turns out that the Feynman "measure" :Px(t) is concentrated on the class of continuous but nowhere differentiable functions, see the remarks on p.18 and in Chap. 2.
1 Introduction
11
the foundations of quantum mechanics  with the notable exception of Pauli, see below. 16 In the orthodox formulation and interpretation of quantum mechanics the idea of an electron's orbit had been completely abandoned by 1925, and thus it appeared that Feynman's path integral approach, which is based in an essential way on the notion of paths, is a regression to improper classical ideas. In a talk given on the occasion of Schwinger's 60th birthday celebration in 1978, Feynman narrated the reaction of Niels Bohr at the famous Pocono conference in 1948 [337]: "That was chaos, and then, all the time I was pushed back, away from the mathematics into my socalled physical ideas until I was driven to the point of describing quantum mechanics as an amplitude for every path, for every trajectory that a particle can take there's an amplitude and Professor Bohr got up and explained to me that already in 1920 they realized that the concept of a path in quantum mechanics  that you could specify the position as a function of time  that was not a legitimate idea and I gave up at this point. As already mentioned, one of the first physicists of the younger generation who immediately appreciated Feynman's approach to quantum mechanics and quantum electrodynamics, was Dyson. Although Dyson [285] did not work with path integrals, he thoroughly understood Feynman's method which permitted him to see the relationships among the conventional operator formulations of quantum electrodynamics, that of Schwinger [838] and Tomonaga [895], and that of Feynman [328330]. One of the most fundamental aspects of path integrals is that they offer a very transparent method to systematically derive the h + 0 of quantum mechanics. To study this limit let us consider the Ddimensional generalization
semiclassicallimit
x(t")=x" K ( x " , x'; t", t ' )  
/
:Dx(t)exp(~R[x(t)])
(1.1.19)
x(t,)=x' of the path integral (1.1.15) where we restrict ourselves to Cartesian coordiD nates, x(t) = (x 1( t ) , . . . , Z D (t)). Here :Dx(t) denotes lIa=l :Dxa (t) and R[x(t)] is the corresponding classical action. At this point the reader is reminded that Dirac [254] had already shown that the multipleintegral construction (1.1.7) contains the quantum analogue of the classical action principle. Feynman remarked [326]: "The points he [Dirac] makes concerning the passage to the classical limit h 4 0 are very beautiful ... ", and he translated Dirac's argument into the path integral language. If h is small, the integrand exp will be a rapidly
((i/li)R[x(t)])
16Here we do not include Bethe, who was most likely the first to judge rightly the value of Feynman's spacetime view, since he started his career in the late 1920s when the principles of quantum mechanics were already invented.
12
Introduction
varying functional of the path x(t), and thus the region in functional space at which x(t) contributes most strongly is that at which the phase of the exponent, i.e., the classical action R, varies least rapidly with x(t) (method of stationary phase). "We see then that the classical path is that for which the integral 1"
R[x(t)] = ]
t tt
dt t
s
(1.1.20)
suffers no firstorder change on varying the path. This is Hamilton's principle and leads directly to the Lagrangian equation of motion." [326] In order to apply the functional analogue of the method of stationary phase 17 to the path integral (1.1.19), we expand the functional R[x(t)] about the classical trajectory XCl(t) in a sort of functional Taylor series
nix(t)] = nCl +
52R[xc,(t)] +
+...
,
(1.1.21)
where Rc, is the classical action evaluated along an actual path xcl (t) of the system, Rcl = Rcl(x", x'; t", t') := R[xcl(t)] 9 (1.1.22) Here xcl(t) is the solution of Hamilton's principle, 5R = 0, with the endpoint conditions x(t') = x', x(t") = x " } s In (1.1.21) 52R[xcl(t)] is a quadratic functional with regard to the quantum fluctuation q(t) = x(t)  x c , ( t ) (q(t') = q(t") = 0)
t" ~ 52s 52R[xcl(t)] = ft'
52s I. Sz"Sxb x=xc, q~(t)qb(t) + 2 5z~6xdx=xc, q~(t)(lb(t) +~
52s
} (l~(t)(lb(t) dt .
(1.1.23)
X=XCI
Inserting the "Taylor series" (1.1.21) into the path integral (1.1.19), we obtain K(x", x'; t", t') q(t")=0
= exp
Rci q(t')=O
Here we have used the fact that the Feynman "measure" transforms as Dx(t) ~ 7)q(t) under the translation x(t) + XCl(t) ~ q(t), since xcl(t) is a fixed function. This is obvious from the lattice definition (1.1.13), since
dx~ =dma(tk ) + d(x~l(tk ) + qa(tk ) ) = dq z. lrSee Sect. 5.2. 1Sin general, there will exist many solutions to the variational problem, see the discussion on p.17 and in Sect. 5.2. Here we ignore this problem.
1 Introduction
13
Retaining only the quadratic functional (1.1.23) in the path integral (1.1.24), the remaining path integration over 7)q(t) can be carried out since it is quadratic in q(t) (Gaussian path integral), and one obtains for small time intervals t"  t ~ the semiclassical (sc) formula 19 K~r (x", x'; t", t')
 (2~rilh)D/2[det(
02Re, ]
i " , x';t",t')] . exp [hnc,(x (1.1.25)
Throughout this book we shall call the formula (1.1.25) the determinant D :=
D(x",x';t",t') := det
Pauli's formula and
02Rc l
Ox~O#b ]
(1.1.26)
the MoretteVan Hove determinant for reasons which will become clear below. Here a~Rcl/ox"ao# b is a D • D matrix (a, b = 1 , . . . , D). We would like to make several remarks: i) It can already be seen from (1.1.24), which is still exact, that the Feynman kernel K can be reduced to a product of two functions, where one of these functions is exactly given by the phase factor exp [(i/h)Rc,] and depends upon the classical path, while the remaining function is the Feynman kernel for a system to proceed from q = 0 at t = t ~ to q = 0 at t = t" and does not therefore depend on x ' , x n, or xcl, being only a function of t ~, t". ii) If the Lagrangian is quadratic to begin with, like in the case of a forced harmonic oscillator, the action functional cannot depend on q more than quadratically, and hence Pauli's formula (1.1.25) is exact. This was already observed by Feynman [326, 330], and a method to compute such Gaussian path integrals was developed by him in his thesis [325]. iii) To the best of our knowledge, the first paper on path integrals, apart from Feynman's, written by a physicist was submitted by C~cile Morette 2~ in 1950 [710]. In this paper we find for the first time the general method of the functional Taylor expansion (1.1.21) applied to path integrals. Pauli's formula (1.1.25) was not directly derived, but rather Morette started from Feynman's ansatz (1.1.12) for the shorttime kernel and determined the normalization factor by a unitarity condition. The formula obtained for this factor was, however, not yet expressed in terms of the determinant 19The exact formula for finite time intervals t "  t ' > 0 is given in equation (5.2.10). 2~ married DeWitt [234236] in 1951 and is identical to MoretteDeWitt [711, 712, 629] and DeWittMorette [147, 237248].
14
Introduction
(1.1.26). It so h a p p e n e d t h a t M o r e t t e , P a u l i a n d Van Hove were simult a n e o u s l y a t t h e I n s t i t u t e for A d v a n c e d S t u d i e s in P r i n c e t o n d u r i n g t h e fall of 1949 a n d the following winter. 21 M o r e t t e discussed her work w i t h Van Hove who i m m e d i a t e l y saw t h e i n t i m a t e r e l a t i o n s h i p to t h e classical HamiltonJacobi theory a n d the calculus of variation (see Sect. 5.2). As a result t h e p r e f a c t o r was (up to a sign factor) f o u n d to have the final f o r m as given in (1.1.25). T h i s result was t h e n included in M o r e t t e ' s p a p e r [710], w i t h an a c k n o w l e d g m e n t to Van Hove, a n d also p u b l i s h e d by Van Hove in [905]. In b o t h p a p e r s the m i n u s sign in t h e d e t e r m i n a n t (1.1.26) is missing, a n d t h u s the semiclassical f o r m u l a for K presented in the two p a p e r s is n o t correct since it is lacking a factor (  1 ) O/2. It s h o u l d be r e m a r k e d t h a t Van Hove's work was not on p a t h integrals. iv) P a u l i was wellknown for h a v i n g been very critical a n d s o m e t i m e s even very h a r s h in s e m i n a r s , say. " P a u l i could be ruthless in d i s m i s s i n g work he c o n s i d e r e d shallow or flimsy: ganz falsch ( u t t e r l y false)  or worse, nicht einmal falsch (not even false)" [396, p.l15]. 22 It is therefore q u i t e 21Historical details can be found in [193]. 22While Feynman was still working at Princeton with Wheeler on their timesymmetric electrodynamics [920], Wheeler asked him to prepare a seminar on that. Feynman remembers [338]: "So it was to be my first technical talk, and Wheeler made arrangements with Eugene Wigner to put it on the regular seminar schedule. "A day or two before the talk I saw Wigner in the hall. 'Feynman', he said, 'I think that work you're doing with Wheeler is very interesting, so I've invited Russel to the seminar.' Henry Norris Russel, the famous, great astronomer of the day, was coming to the lecture! "Wigner went on. 'I think Professor von Neumann could also be interested.' Johnny von Neumann was the greatest mathematician around. 'And Professor Pauli is visiting from Switzerland, it so happens, so I've invited Professor Pauli to come'  Pauli was a very famous physicist  and by this time I'm turning yellow. Finally, Wigner said, 'Professor Einstein only rarely comes to our weekly seminars, but your work is so interesting that I've invited him specially, so he's coming, too.' " . . . "Then the time came to give the talk, and here are these monster minds in front of me, waiting!" . . . "But then a miracle occurred, as it has occurred again and again in my life, and it's very lucky for me: the moment I start to think about the physics, and have to concentrate on what I'm explaining, nothing else occupies my mind  I'm completely immune to being nervous. So after I started to go, I just didn't know who was in the room. I was only explaining this idea, t h a t ' s all. "But then the end of the seminar came, and it was time for questions. First off, Pauli, who was sitting next to Einstein, gets up and says, 'I do not sink dis teory can be right, because of dis, and dis, and dis,' and he turns to Einstein and says, 'Don't you agree, Professor Einstein?' Einstein says, 'Nooooooooooooo,' a nice, Germansounding 'No,'  very polite. 'I find only that it would be very difficult to make a corresponding theory for gravitational interaction.' . . . "I wish l had remembered what Pauli said, because I discovered years later
1 Introduction
15
remarkable that Pauli was, to the best of our knowledge, the first among the physicists of the older generation, having laid the foundations of quantum mechanics, who fully appreciated the new approach developed by Feynman. From a letter [766, Letter no. 997] dated January 8, 1949, which Pauli sent from Z/irich to Dyson, we can quite precisely infer when it happened that Pauli got interested in Feynman's approach to quantum electrodynamics. In this letter Pauli writes: "I thank you very much for sending your paper. It was not easy to read for us because the 'Feynman theory', which you compare with the SchwingerTomonaga formalism was entirely unknown here and we had to reconstruct it from your paper. Obviously, Pauli refers to Dyson's first paper [285] 23 which was received on October 6, 1948 by The Physical Review and which was sent to him personally by Dyson. On May 10, 1949 Pauli's famous paper with Villars [768] was received by Reviews o/Modern Physics, which contains what is nowadays known as the PauliVillars regularization, and in this paper the two papers by Dyson [285], and Feynman's talk at the Pocono Conference as well as his paper [327] are cited. Thus we can almost be sure that Pauli had read the three papers [326329] of Feynman's when he arrived in Princeton on November 29, 1949, where he stayed until the end of April 1950. 24 During his stay in Princeton Morette and Van Hove presented to Pauli at the occasion of an appointment with him the semiclassical formula (1.1.25). As a result of the discussion with Morette and Van Hove, Pauli wrote a couple of research notes entitled Feynman's Methode der Lagrangefunktion (PN 8/121123), 25 Van Hove (PN 8/150) and Diskutiere Van Hove's Formel (PN 8/154159). In these notes Pauli corrected the sign factor mentioned under iii) and then considered the semiclassical formula (1.1.25) as an ansatz for small but finite time intervals t"  t'. After some calculations, which are very similar to those worked out by him in his article for the Handbuch der Physik published in 1933 [764], he obtained the important (exact) result that Ksr (1.1.25), satisfies the Schrbdinger equation up to terms of order h 2, called "wrong n  1 / 2 wV2x , ,n~l / 2 , the coefficient of order h ~ beterms", proportional to ~ ing the HamiltonJacobi equation, and that of order h 1 the continuity equation satisfied by the probability density D ( x " , x'; t", t'), the square of the amplitude of Ksc. that the theory was not satisfactory when it came to making the quantum theory. It's possible that that great man noticed the difficulty immediately and explained it to me in the question ... " 23The second one was received only on February 24, 1949 by The Physical Review. 24The dates are taken from [766, p. 711 and p. 915]. 25These notes are in the Pauli Archives at CERN, Geneva. The meaning of, e.g., PN 8/121 is "Pauli Nachlass" Box 8, p. 121.
16
Introduction
During the winter semester 195051 Pauli gave a course at the ETH Ziirich on "Ausgew~ihlte Kapitel aus der Feldquantisierung" [765]. The lecture notes contain an Appendix entitled "Der Feynman'sche Zugang zur Quantenelektrodynamik" (Feynman's approach to quantum electrodynamics). There one finds in equation (172) precisely the semiclassical formula (1.1.25) and the proof that IQc satisfies the SchrSdinger equation up to terms of order h 2. Furthermore, Pauli shows that if IQc is inserted for the shorttime kernel in the Ddimensional generalization of the multipleintegral representation (1.1.11), one obtains the exact Feynman kernel. From reading the lecture notes it becomes clear that Pauli understood and appreciated Feynman's path integral approach completely. However, it is interesting to observe that he did not quote Van Hove nor Morette. It seems [193] that one of the reasons why Pauli did not react to Morette's functional approach is that at that time it was not known that the saddle point approximation to Feynman's path integral yields the same result as the timedependent WKB approximation. One can almost be sure that Pauli's opinion at that time is adequately expressed in a comment which was made several years later by Gel'fand and Yaglom in their famous review [376]26 on integration in functional spaces. Commenting in footnote 21 on Morette's paper [710], they wrote: "We note, however, that the strictness of the quoted proof is substantially lowered due to the fact that the question of the precise meaning of functional integrals studied was not discussed." Since Pauli had checked directly that Ks~, (1.1.25), satisfies the SchrSdinger equation up to terms of order h2, there was no doubt about his proof. This explains Pauli's remark at the end of a letter of April 1951 [767, letter no. 1230], congratulating Bryce SeligmanDeWitt and
[email protected] Morette on their marriage: ~7 "By the way,
[email protected] may be interested in the way I have treated the Feynmanaction principle in my mimeographed lectures. It is a kind of generalization of the WBK method to timedependent solutions." v) We do not know of any other papers written by Pauli in which he treats Feynman path integrals. There is, however, another clear indication showing that Pauli considered Feynman's Lagrangian approach to quantum mechanics as an important alternative to the conventional operator approach: in the fall of 1951 Pauli accepted Choquard as a Ph D student and asked him to study the higher order terms in the semiclassical expansion of the Feynman kernel for small but finite time intervals, in particular for Lagrangians which depend on x more than quadratically. Choquard received his Ph D in December 1953 and published his thesis in Helvetica Physica Acta [192]. Choquard's paper contains a very thorough and systematic study of the semiclassical approximation to the Feynman kernel K generalizing Pauli's formula (1.1.25). As 26In the following we quote from the English translation published in 1960. 27See also footnote 20.
1 Introduction
17
specific examples he considered twodimensional quantum billiards (La "boule de billard") and confinement potentials (i.e., anharmonic oscillators with V ( x ) .. x 2k, k > 1), which in recent years play an important r61e in various fields of modern physics. Among several new results which he obtained, he made the important observation that for such systems Pauli's formula has to be modified in an essential way by replacing it by an infinite sum of the form (T := t"  t' > 0) 28
Ksr
1
~
x/2 iR./h
(27rib)D/2 .. D ,
e
(1.1.27)
n=0
Here the nth term with action P~ is the contribution of the nth member of an infinity of classical trajectories passing through x' at time t ~ and x " at time t" for given x' and x " and fixed time interval T. This reflects the important fact that, since the time T is fixed, but not the energy of the classical paths, there exist infinitely many solutions to Hamilton's principle, ~R = 0, where Rn denotes the classical action evaluated along the nth path. Furthermore, he could show that there exits a minimal time tm such that his semiclassical formula (1.1.27) holds for 0 < T < tin. The time tm is determined by the socalled conjugate points of the classical trajectories (in the sense of Jacobi), which are the points at which the MoretteVan Hove determinants Dn become singular. ~9 The singularities of Dn have been investigated, in the context of semiclassical quantum mechanics, for the first time by Choquard [192]. vi) The timeevolution kernel K is the primary object in Feynman's path integral approach and contains the complete information about a given quantum mechanical system, i.e., wave functions and energy levels. But in order to extract this information from K, one needs to have a (semiclassical) formula for it which is valid for finite times, i.e., beyond the conjugate points. 3~ This is a difficult problem and requires among other things nontrivial results from HamiltonJacobi theory and the calculus of variations in the large. Going beyond the conjugate points was first achieved by Gutzwiller 31 in 1967 [479] who found the correct generalization of Choquard's formula (1.1.27) valid for arbitrary times T. Gutzwiller made this formula as the starting point for the derivation of the by now famous Gutzwiller trace formula, which is the basic semiclas2SThe systems considered in [192] are timeindependent and therefore K depends on T only, i.e., K(2:", z'; t", t') = I((~", ~'; t"  t', O) =: I((~", x'; T). 29See Sect. 5.2 for a defmition of conjugate points and their geometrical meaning. 3~ example, if one wants to calculate the Green function G, see equation (2.1.25), one has to integrate K over the whole time interval, T E (0, oo). 31It is, presumably, not by chance that Gutzwiller is also a former student of Pauli. He wrote his diploma thesis (on meson theory and the anomalous magnetic moment of the proton!) under the direction of Pauli in 1949.
18
Introduction
sical quantization rule for strongly chaotic systems [483, 869]. This will be discussed in Chap. 5. In answering the question of whether the new (third) formulation of quantum mechanics had a favourable reception, we have so far paid attention only to the physicists. Since the path integral is, after all, a mathematical object, it is interesting to enquire about the reaction among the mathematicians. It is quite remarkable to learn that there appeared already in 1949 an interesting paper by the mathematician Mark Kac [555] which was written, as the author indicates, under the strong influence of Feynman's work. 32 Kac had worked in probability theory [554], in particular on the extension of Wiener's work [930] on Brownian motion. In this work there had already appeared a special measure in the space of continuous functions, called Wiener measure. Kac realized that if the path integral (1.1.15) is analytically continued to purely imaginary time ("Euclidean time"), t +  i t , see footnote 3, it can be rewritten in terms of the welldefined conditional Wiener measure. In fact, the Feynman path integral can then be interpreted as the mean value (expectation value) of the real functional exp [  ~ ftt,'' V(x(t))dt] over the trajectories of a Brownian particle, also called a diffusion or Wiener process. Thus Kac was able to show that Feynman's path integral, considered in Euclidean time, is a welldefined functional integral. Following Kac's article a lot of papers appeared in the mathematical literature developing these same ideas further. 33 In quantum mechanics, the reformulation of Feynman's path integral expression for the kernel K in terms of the Wiener measure is well known today as the FeynmanKac Formula, see e.g. [397, 706,854]. Kac later felt that he was better known as the K in FK than for anything else in his career [556, p.115116]. Although we do not know whether Feynman was aware of Wiener's work in probability theory, it is quite clear that he had realized the stochastic nature of the dominant paths in his path integral. Already in 1942 Feynman wrote in his Ph D thesis [325]: "Although the average value of the displacement of a particle in the timedt is vdt, where v is the mean velocity, the mean value of the square of this displacement is not of order dt ~, but only of order dt." And in his 1948 paper he even refers to Brownian motion: "The 'velocities' (xj_ 1  Xj)/r which are important are very high, being of order ( h / m e ) 1/2 which diverges as e + 0. The paths involved are, therefore, continuous but possess no derivative. They are of a type familiar from study of Brownian motion." [326]. That Feynman was "familiar", indeed, with the theory of Brownian motion and, more generally, with the theory of diffusion processes is well known. During the war Feynman worked at Los Alamos in the theoretical 32Kac heard Feynman describe his path integral at Cornell, see Gleick [396, p. 249]. 33For a rather complete mathematical review on Integration in Functional Spaces and its Applications in Quantum Physics, covering the years until 1955, the reader should consult the famous paper by Gel'fund and Yaglom [376].
1 Introduction
19
division, and in 1944 Bethe, who was in charge of this division, decided to make Feynman a group leader. The official name of the group was T4, Diffusion Problems [396, p. 171]! We also know that Feynman was during these years very close to the great mathematician John von Neumann who served as a travelling consultant and helped Feynman and his group with the numerical computations on the first computers available then. It is hard to believe that von Neumann did not tell Feynman about Wiener's work, knowing that Feynman was working on diffusion problems! Coming back to the FeynmanKac formula, we would like to make another remark. By reading the papers which appeared on this subject during the last fifty years, we cannot avoid getting the impression that some authors consider Feynman's original work on the path integral as a minor contribution relative to the rigorous work of Kac and other mathematicians. We will not comment on this, but rather cite Kac [556] himself who certainly knew how to judge Feynman's contribution: "There are two kinds of geniuses, the 'ordinary' and the 'magicians'. An ordinary genius is a fellow that you and I would be just as good as, if we were only many times better. There is no mystery as to how his mind works. Once we understand what they have done, we feel certain that we, too, could have done it. It is different with the magicians. They are, to use mathematical jargon, in the orthogonal complement of where we are and the working of their minds is for all intents and purposes incomprehensible. Even after we understand what they have done, the process by which they have done it is completely dark. They seldom, if ever, have students because they cannot be emulated and it must be terribly frustrating for a brilliant young mind to cope with the mysterious ways in which the magician's mind works. Richard Feynman is a magician of the highest caliber." Our intention in this Introduction was to give the reader some historical insights into the remarkable development of Feynman's path integral approach, and to enable him or her to see things in their right perspective. Since some of the facts which we have touched upon are not so well known, we thought it would be worthwhile to present them here. The decades since the early 1950s, have seen a triumphal success of Feynman's path integral method. The applications cover many different areas, notably in physics, chemistry and mathematics. In this Handbook we shall mention a n d / o r list a large number of these applications, i.e., a large number of path integrals together with an extensive list of almost 1000 references. Our book is organized as follows. In Chap. 2 we give an introduction to the General Theory of Path Integrals. This chapter is selfcontained and is written for those readers who have not yet studied path integrals, but have a good knowledge of the fundamentals of quantum mechanics as covered by standard courses in theoretical physics. Sections 2.1 and 2.2 contain the basic definitions and properties of path integrals, while Sections 2.3 and 2.4 provide some rules for how to compute simple path integrals. Sections 2.52.11 are written on a more advanced level, with an increasing degree of difficulty. The
20
Introduction
techniques described there have been developed only recently, and it is only with these new techniques that, for example, it has been possible to compute the path integral for the hydrogen atom, which is the prototype example of quantum mechanics, see Sect. 2.10. In Chap. 3 we discuss and compute in detail the path integrals which we have called basic path integrals. It turns out that practically all path integrals, that can be calculated in closed form, can be in some way or another reduced to these basic path integrals. Chapter 4 contains an introduction to perturbation theory, which is most elegantly derived from the path integral. This is the quantum mechanical analogue of Feynman's original derivation of the Feynman rules in quantum electrodynamics. However, in the nonrelativistic case, treated in Chap. 4, we do not rephrase the formulae in terms of graphs (although this is possible). In Sect. 4.2 we discuss an example for which the perturbation series can be summed up exactly. Section 4.3 deals with an application of path integrals to statistical mechanics, in particular with the partition function and the socalled effective potentials. In Sect. 4.4 we discuss the semiclassical expansion of the path integral about the harmonic approximation. In Chap. 5 we give a short introduction to the semiclassical theory and its recent applications in the field of quantum chaos. The basic formula is Gutzwiller's expression for the Feynman kernel which is derived in Sect. 5.2. This completes the work started by Feynman, Morette, Van Hove, Pauli and Choquard, as described in the foregoing introduction. In Sect. 5.3 we derive the corresponding semiclassical formula for the Green function and, finally, Sect. 5.4 contains a discussion of the Gutzwiller trace formula which is the basic relation in the theory of quantum chaos. Our final Chap. 6, which makes up the largest part of this Handbook, contains a rather complete table of path integrals in nonrelativistic quantum mechanics, including supersymmetric quantum mechanics, and statistical mechanics. The path integrals in this table are classified according to our basic path integrals introduced in Chap. 3. 34 A comparison of the table with the known exact solutions of the SchrSdinger equation shows that it is possible nowadays, with the modern techniques described in Chap. 2, to solve all path integrals for which the SchrSdinger equation can be solved. To each path integral listed in the table we attach a comprehensive list of references which provides for the reader easy access to the original literature and thus offers the possibility of having a closer look at the derivation of the various path integrals and their applications in different fields. For the sake of completeness we include some references corresponding to relativistic path integral solutions, i.e., for the KleinGordon [70, 82, 117, 118,212,226,324, 331,394,510, 613,653,654, 810,828,912] and the Dirac equation [17, 24, 55, 98, 117, 118i 203,212,226,340,243,372,370,393,508, 34A brief outline of our classification of path integrals was presented in our previous papers [469, 470].
1 Introduction
21
536538,718,731,738,754, 769,789,796,803,806,879], which are, however, incomplete; we do not dwell on the mathematical definitions, problems and ambiguities of these path integral representations. For padic path integrals, see e.g. [154, 693, 759, 846,907, 944]. For the interested reader who wishes to study certain fields or applications in more detail, we give the following list of textbooks on Feynman path integrals: Feynman and Hibbs Quantum Mechanics and Path Integrals [340] Feynman Statistical Mechanics [334] Schulman Techniques and Applications of Path Integration [828] Simon Functional Integration and Quantum Physics [854] Glimm and Jaffe Quantum Physics: A Functional Point of View [397] Albeverio and HceghKrohn Mathematical Theory of Feynman Path Integrals [18] Antoine and Tirapegui Functional Integration: Theory and Applications [27] Dittrich and Reuter Classical and Quantum Dynamics [257] Exner Open Quantum Systems and Feynman Integrals [306] Inomata, Kuratsuji and Gerry Path Integrals and Coherent States of SU(2) and SU(1, 1) [528] Junker Supersymmetric Methods in Quantum and Statistical Physics
[5501 Kac, Uhlenbeck, Hibbs and van der Pol Probability and Related Topics in Physical Sciences [557] Khandekar, Lawande and Bhagwat PathIntegral Methods and Their Applications [587] Kleinert Path Integrals in Quantum Mechanics, Statistics and Polymer Physics [613] Langouche, Roekaerts and Tirapegui Functional Integration and Semiclassical Expansions [637] Roepstorff Path Integral Approach to Quantum Physics [801] Smolyanov and Shavgulidze Continual Integrals [857] Weiss Quantum Dissipative Systems [915] Wiegel Introduction to PathIntegral Methods in Physics and Polymer Science [927]. In addition to these textbooks the reader may consult the Special Issue on Functional Integration in Journal of Mathematical Physics edited by DeWittMorette [240], and the following conference proceedings: Arthurs Functional Integration and Its Applications [35] Papadopoulos and Devreese Path Integrals and Their Applications in Quantum, Statistical, and Solid State Physics [755] Swanson Path Integrals and Quantum Processes [884]
22
Introduction The proceedings of the conference series Path Integrals from meV to MeV beginning in 1985 [484, 662, 819, 153, 414,817, 940].
Although, in this book, we shall not discuss the application of path integrals to quantum field theory, we would like to give the following list of textbooks in which these matters are discussed: Becher, BShm and Joos Gauge Theories of Strong and Electroweak Interactions [66] Bogoliubov and Shirkov Introduction to the Theory of Quantized Fields [107] Creutz Quarks, Gluons and Lattices [210] Das Field Theory. A Path Integral Approach [217] Faddeev Introduction to Functional Methods [311] Faddeev and Slavnov Gauge Fields: Introduction to Quantum Theory [313] Feynman Quantum Electrodynamics [335] Itzykson and Drouffe Statistical Field Theory [533] Itzykson and Zuber Quantum Field Theory [534] Kugo Eichtheorie [626] Lee Particle Physics and Introduction to Field Theory [645] Montvay and Miinster Quantum Fields on a Lattice [707] Popov Functional Integrals in Quantum Field Theory and Statistical Physics [782] Ramond Field Theory [790] Rebbi Lattice Gauge Theories and Monte Carlo Simulations [793] Rivers Path Integral Methods in Quantum Field Theory [799] Roepstorff Path Integral Approach to Quantum Physics [801] Rothe Lattice Gauge Theories [808] Swanson Path Integrals and Quantum Processes [884] Weinberg The Quantum Theory of Fields [916]. Several of Feynman's original papers [326, 328330] as well as Dirac's paper [254], and some other seminal papers on quantum electrodynamics are reprinted in Quantum Electrodynamics [840] edited by Schwinger.
2 General
Theory
2.1 T h e F e y n m a n K e r n e l a n d t h e G r e e n F u n c t i o n
Let us start with the simplest case, i.e., with the onedimensional motion of a particle of mass m under the influence of the timeindependent force F(x) =  d V ( x ) / d z , where V(x) denotes the potential, and x = z(t) C IR the classical trajectory as a function of time t E IR. The classical dynamics can be completely formulated in terms of the classical Lagrangian s163
V(x)
(2.1.1)
or, equivalently, in terms of the classical Hamiltonian H = H(p,x) : = p ~  s
1 ~m p + V(x) .
(2.1.2)
Here /c = dx(t)/dt is the velocity, and p := 0s = rode the generalized momentum conjugate to x. The standard formulation of quantum mechanics starts from the Hamiltonian (2.1.2). Working in the Schr5dinger picture, the canonical variables (z, p) are replaced by the timeindependent Hermitian operators (x, p) which act on timedependent state vectors [gr(t)) E 7 / o f a separable Hilbert space 7/. The algebra of the operators is fixed by the Heisenberg commutation relation [x_,p] := xp  p_x = i h , (2.1.3) where h denotes Planck's constant divided by 27r. At a given time t, the physical state of the quantum mechanical system is completely described by the vector Igt(t)) E 7/. Replacing the canonical variables in the classical Hamiltonian (2.1.2) by the corresponding operators leads to the welldefined quantum Hamiltonian (operator) I:I := H(p,x) = 2~p 2 + V(x) .
(2.1.4)
(In the general case one encounters operatorordering problems which will be discussed in subsequent sections.) The quantum mechanical timeevolution is governed by the SchrSdinger equation i h d l ~ ( t ) ) = Ul~(t)) . (lg
(2.1.5)
24
General Theory
Knowing the state [~(t')) E ~ / a t the initial time t', the problem of quantum mechanics consists in computing the state of the system at an arbitrary final time t" > t ~. The general solution of (2.1.5) can be written as [~(t")) = U(t", t')[$'(t')) ,
(2.1.6)
where U denotes the unitary timeevolution operator satisfying the operator equation
i h~&zU(t",t' ) = HU(t",t')
(2.1.7) Ot with the initiM condition U(t ~, t ~) = 1. For the timeindependent Hamiltonian (2.1.4) one immediately obtains the explicit solution (T := t"  t')
U(t",t') = exp (  ~H_ T)
(2.1.8)
which fulfils the composition law U(t",t') = U(t",t)U(t,t')
(2.1.9)
for arbitrary times t ~, t, t". In almost all practical calculations one does not work in the abstract Hilbert space, but rather in the socalled zrepresentation, respectively coordinate space representation. Consider the eigenvectors Ix) of the position operator ~ satisfying x[x)  xlz) (2.1.10) with the continuous spectrum x E IR. (Here we restrict ourselves to systems where the onedimensional motion of the particle takes place on the whole real line without additional topological constraints. Systems where the motion is confined to smaller regions, e.g., the halfspace x > 0, will be discussed in later sections.) Then we have the (continuum) normalization
(x'lx) = 6(x'  x)
(2.1.11)
and the completeness relation
/dz
Iz)(~l = 11
I
(2.1.12)
5(z'  x) denotes the Dira~ deltafunction. Using (2.1.12) we get the following transformation formula from the abstract Hilbert space 7/ to the xrepresentation I,(t)> 
f dxI x ) ( x l r
=
fRdzr162
(2.1.13)
where ~P(z, t) denotes the (complexvalued) Schr6dinger wave function corresponding to the state vector I~(t)) defined by
2.1 The Feynman Kernel and the Green Function
25
# ( x , t ) := (x]#(t)) = (#(t)lx)* .
(2.1.14)
With the help of (2.1.11) one immediately derives from (2.1.13) the normalization
(#(t)l#(t)) =
J dr 1~(~, t)] 2 =
1
(2.1.15)
which shows that ~(x, t) E s In the xrepresentation there exists an explicit realization of the Hermitian operators (_x,p) satisfying the commutation relation (2.1.3): _x acts on the wave function ~t as a multiplication operator, while p acts as the differential operator  i h O / O x . Then the Schrhdinger equation (2.1.5) takes the standard form
i h O~(x' Ot t) _ H_Jt(x,t)
(2.1.16)
with the differential operator (Schrhdinger operator) tJ, defined by
H~ 
h2 02 2m Ox 5 + V(x) .
(2.1.17)
Using the definition (2.1.14) and the relation (2.1.12), one immediately derives the timeevolution equation for the Schrhdinger wave function (t" > t ~)
#(x",t") =/dx'g(x",x';t",t')#(z',t')
(2.1.18)
with the (retarded) Feynman kernel (T = t"  t') K(x", x'; t", t') := (x"lU(t", t')laz')O(t ' '  t') =<x"]exp(~_HxT
) x'}O(T)
.
(2.1.19)
Here O(T) denotes the Heaviside step function defined by 1
O(T):=
1 0
T _> 0 T 0.
26
General Theory
but is homogeneous in time, see (2.1.19). We therefore usually write in the timeindependent case
K ( x " , x ' ; T ) := K ( x " , x ' ; t " , t ' ) = K(x",x';T,O)
.
(2.1.21)
Using O~(T) = 5(T) it is easy to see that K satisfies the inhomogeneous Schr6dinger equation
(0~h~g~,, )I((x",x';T)=ihS(x"x')5(T)
(2.1.22)
with the initial condition (see (2.1.11) and (2.1.19)) lim K(x", x'; T) = 5(x"  x') .
(2.1.23)
T~O+
The composition law (2.1.9) implies for the kernel K (t' < tz < t")
K(x",
=f
IC(xt',Xl;t't,tl)I((Xl,X';tl,t t)
(2.1.24)
which is an important law for the composition of amplitudes for events which occur successively in time [326, 340]. There are many quantum mechanical systems for which the timedependent kernel K cannot be given in explicit form, but instead its Fourier transform with respect to time can be explicitly written down. We are thus led to define the energydependent (outgoing) Green function
G(z", x'; E) :=
i
dTei(E+i
e)T/h K(x,t,
x'; T)
,
(2.1.25)
where a small positive imaginary part (r > 0) has been added to the energy E. 2 From (2.1.19) we obtain
G(x",x';E)=<x"
_H~  E  i e l
x'> .
(2.1.26)
In mathematics, the operator ( A  z ) z, z E C\spec(_A), is called the resolvent of a given operator A, and thus G is the resolvent kernel of the Hamiltonian _H~ in the coordinate representation. Knowing the Green function, we can recover the Feynman kernel via the inverse Fourier transform
K(x", x'; T) = / r t 27riedE_ i ET/h G(x", zt; E) .
(2.1.27)
The Green function satisfies the inhomogeneous SchrSdinger equation
(H~,,  E)G(x", x' ; E) = 5(x"  x') .
(2.1.28)
Usually we shah not explicitly write the i e, but tacitly assume that the various expressions are regularized according to this rule.
2.1 The Feynman Kernel and the Green Function
27
In the generic case, the kernels K and G will decompose into two terms corresponding to the contributions from the bound states (discrete spectrum) and the scattering states (continuous spectrum) of a given quantum system. Let us briefly discuss the simplest case for which both kernels can be explicitly given, i.e., the free particle (V(x)  0) which is described by the
free Hamiltonian
1 2 _Ho := ~~mp
(2.1.29)
9
Since I:I0 depends on p only, it is natural to go to the prepresentation and consider the eigenvectors IP) of the momentum operator p
plp)=plp),
/rtdplp)(pln
(p'lp)=5(p'p),
(2.1.30)
with p E IR. Then the most general solution of the free SchrSdinger equation reads ("wave packet") Ik~(t)) = f ~ dp4i(p) eiE(p)t/n IP) with ~(p) C s
(2.1.31)
and
HolP)=
E(p) =
E ( p ) lp) ,
p2 ~
.
(2.1.32)
Obviously, in this case the energy spectrum is continuous, E(p) >_O, and the corresponding wave functions are plane waves (free scattering solutions) ~p(z) := (zip)  ~ ~/21rh
satisfying the
eipz/h
(2.1.33)
orthogonality relation J dz
= J(p'  p)
o
(2.1.34)
(Notice that while the plane wave (2.1.33) is not normalizable, the wave packet (2.1.31) is square integrable since ~(p) E s which implies for the corresponding wave function ~(x, t) = (xl~(t)) E s One then obtains for the free Feynman kernel
Ir
exp (  h H o T )
=/~dp" j~ dp' (x",p") 0 and define x0 := z', to := t ~, X N := X tt, tN :~ ttt, i.e., tj  t' +ej, x(tj) = xj, j = 0, 1 , . . . , N . Using the definition (2.1.21) we can rewrite (2.2.1) as N1
N1
K(x",x';T)= H /R dzk H K(Xj+l,Xj;e) . k=l
(2.2.2)
j=0
This relation does not seem to be very useful since the unknown kernel K occurs on both sides. The crucial point is to consider the limit of an infinitesimally fine lattice, i.e., N ~ 0% T fixed, which is equivalent to the limit r + 0. In this limit the r.h.s, of (2.2.2) depends only on the shorttime kernel
K(Xj+l,Xj;e)(Xj+l exp (  ~ e _ H ) x j ) = (xj+llexp (  h cHo  ~cV(x)) xj )
(2.2.3)
which can be exactly calculated up to terms of O(e2). With the help of the Zassenhaus formula [717, 881,932] exp [e(A + _B)] = exp (e_A) exp (eB) xexp one obtains
( ~[A,B])exp
[_A,_B]]+ I A [_A,B_]]}
(2.2.4)
32
General Theory
K(zj+,,zj;e)= O)
34
GeneralTheory _
( m ~N/2Nlf
= lim
k=l [ 1N1( m
• exp
 ~ "~:o ~ ( * ~ + '  *~)~ + ~V(.~)
.(t"):."f [ J DEx(t) exp ~(t')=x' =:
J
.xp
m'2
~
~x
(/')  g1
~w[~]exp
, v(~(t)) d~
)1
+ V(z)) ]dt ,
where DW[x] denotes integration with respect measure [397]. In the mathematical literature for the Euclidean Feynman kernel KE(T) is formula [554, 794, 854], and instead of path
(2.2.9) to the conditional Wiener the representation (2.2.9) called the FegnmanKac integration one speaks of
.functional integration.
In the path integral (2.2.9) the contribution from a given path x(t) is proportional to exp(RE[z(t)]/h) and thus positive definite, where RE denotes the Euclidean action tH
RE[x(t)] :=
/
s
.It t
A comparison of the Euclidean action (2.2.10) with the standard action (2.2.8) shows that the Euclidean path integral can be interpreted as describing a particle moving in a potential minus V, i.e., in the inverted potential V(x). This observation lies at the heart of the socalled instanton approximation, see e.g. [200,638], which gives in the semiclassical limit (h ~ 0) the dominant nonperturbative contribution to the amplitude for transmission through a potential barrier (barrier penetration is not seen in any order of perturbation theory in h). vii) If the time T in the Euclidean path integral (2.2.9) is redefined to be ~h, one obtains the path integral formulation of the density matrix p in statistical mechanics [334,340]
p(=",
~(~n)=~"
=
I='>
J ~Ex(s)exp[~o"(~'~(.)+V(x(s)))ds] i2.2.11 ) x(o)=~,
2.2 The Path Integral in Cartesian Coordinates
35
with /3 = (kB 9 temperature) 1, and kB Boltzmann's constant. From (2.2.11) one easily derives a very powerful representation for the partition function Z (or the free energy F) in statistical mechanics by taking the trace oo
Z=e 0F=Tre zH=Ee~B"
=/dxp(x,x;/3)
am0
~(0)=~
(2'2"12/
Here we have assumed that _Hhas a purely discrete spectrum, see (2.1.41), otherwise the contribution of the continuous spectrum must be properly treated. viii) The potential V(x) appearing in the action may also be complex valued. The imaginary part of the potential can be understood as a source, respectively a sink, for particles [936]. A complex potential can also appear from a transformation of a timeindependent Hamiltonian to a timedependent one [440,773], which has the consequence that the new Hamiltonian does not conserve the energy, which is exactly balanced by the imaginary part of the potential to guarantee energy conservation of the entire system. The corresponding term can also be interpreted as a "pathdependent measure" [773]. ix) The derivation of the path integral (2.2.6) can be put on a rigorous mathematical basis by starting not from the composition law (2.2.1) but instead from the definition (2.1.19) K ( x " , z ' ; T ) = ( x " exp [  ~ T ( _ H 0 + V(_x))] x'}O(T)
(2.2.13)
and employing Trotter's formula [569,854,896]: Let _A and B_ be selfadjoint operators on a separable Hilbert space so that _A+ B is selfadjoint. Then exp[it(A +/3)] = N+~slim[ e x p ( i t A / N ) e x p ( i t B / N ) ]
N
(2.2.14)
If furthermore, d and B are bounded from below, then (t > O) exp[t(_A + _B)] = slim [exp(t_A/N)exp(t_B/N)] g .
(2.2.15)
N~eo
x)
If the Hamiltonian H is time dependent, the solution (2.1.8) for the timeevolution operator has to be replaced by the FegnmanDyson formula
U(t",t') = Texp
 ~
IJ(t)dt
(2.2.16)
36
General Theory := ll~i t"dtlH_(ti)+(h)2ftt"dtlfttadt2H_(tl)H_(t2)+jt, f
"'" (2.2.17)
and the Feynman kernel is, in an obvious generalization of (2.1.19), defined by
(2.2.18) (Here T denotes the timeordered product.) In this case the lattice definition of the Feynman path integral reads
K(z", z'; t",t') = lim

( m )~/~N1s
9:,,, [,,,(i I ,, 9
(2.5.24)
In particular, the coherent state with z = 0 is identical to the Fock vacuum
10>. Summarizing, we observe that the operator a acts on coherent states by multiplication by z, while at is represented by O/Oz. Furthermore, we have the completeness relation I  11
cd,(z)lz)(z
(2.5.25)
and the scalar product (zlz') = e ~'~'
(2.5.26)
From (2.5.25) and (2.5.26) it follows that (zlz') acts in the Bargmann representation like a Dirac delta distribution (reproducing kernel) r
= (zlr =
= / dl~(z')(zJz')(z'lr I
*
t
du(z ) e ~ ~ r
(2.5.27)
while the operators a t and a act as (atr (ar
: : (zlatlr = z*r := (z[a[r
= ~~r
,
(2.5.28) .
(2.5.29)
An arbitrary operator A = A(a,_a t) in Fock space can be written as oo
_A= ~
]m)Am'*(nl ,
Am,, := (ml_A[n) .
(2.5.30)
rr$,,,=0
Its matrix representation is given by oo
(zl_AIz') = y~ f,,,(z*)A,n,,f,,(z') r t l , , , ~0
=: A(z*, z') ,
(2.5.31)
where the analytical function A(z*, z') plays the r61e of an integral kernel which represents the action of the operator A on a state r in the Bargmann representation
48
General Theory (Ar
= f dp(z')A(z*, z')r
.
(2.5.32)
The product of two operators A and B_ possesses the matrix representation
(_A0)(z*, z')  (zl_A_BIz') f,
J dlt(z")A(z*, z")B(z"*, z') .
(2.5.33)
Using the commutation relation (2.5.1) it is always possible to bring an operator _Ainto normal ordered form with all the operators _a~"standing on the left of the operators _a, i.e., co
A= ~
Ckt(at)k(a) l .
(2.5.34)
k,l=O
(This should not be confused with "normal ordering" or Wick's ordering denoted by a doubledot symbol, e.g., : _a_at := _at_a.The normal ordered form of the operator A = _a_at is A = ata + 11 #: A :.) The normal symbol of the operator (2.5.34) denoted by AN(z *, z') is defined by [311,313] co
aN(z*,z ') := ~
ck,z*kz ''
(2.5.35)
k,l=O
The relation between the kernel (2.5.31) and its corresponding normal symbol A N is given by oo
A(z*, z') = (z]A]z') = ~
ck, z*kz"(z[z ')
k,l=O
= eZ'Z'AN(z*, z') .
(2.5.36)
Thus to obtain the matrix representation of an arbitrary operator _Ain the Bargmann representation, one first brings A into its normal ordered form (2.5.34), then forms the normal symbol (2.5.35) and finally just multiplies Aiv by e~'z'. For more details on coherent states, the reader is referred to Faddeev [311], Faddeev and Slavnov [313], and Klauder und Skagerstam [602]. 2.5.2 T h e P a t h Integral. For the onedimensional Feynman kernel we have
K(z", z'; t", t') = <x"lU(t", t')Iz')e(t"  t') = f d/~(z") f
dl~(z')(z"lz")U(z"*,z';t",t')O(ff' t')(z'lz')
oo
= ~
r
(2.5.37)
2.5 Path Integration Over Coherent States
49
with the timeevolution kernel
U(z"*, z'; 1", t') := (z"lU(t" , t')lz' >
(2.5.38)
and the transitionmatrix element [340, p.144]
Km,~(t",t') := (mlU(t",t')ln)~9(t"  t')
=f =f
d.(z')/m(z"lV(z"',z';t",t'lO(t"l')/.(z"l (2.5.39)
Here we have used ( r
(xln)) oo
< lz> =
(2.5.40) n0
For t H > t ~ we have the expansion
g(z"*,z';t",t')=
fi
fm(Z"*)Kmn(t",t')fn(z'),
(2.5.41)
rr~,n:O
which shows that the timeevolution kernel U(z"*, z'; t", t') is the generating function of the transition amplitudes K,,m
K,,,(1",1') = ~ 1
0 m+n U ( z t'* , z'; t H, t') O(z,,.)~az,,,
(2.5.42) zll l*:zl:O
Notice that IK~,, (1", t')12 is the probability for the transition during the time interval 1"  1' > 0 from the initial state In) at time t' to the final state Ira) at time 1~. In particular
Koo(t", t') = U(z"*, z'; 1", t') z ...._z,o_
(2.5.43)
is the vacuumvacuum transition amplitude. From (2.5.39) we derive lim,,,~t, Kmn(t", t') = ~,nn, and thus obtain from (2.5.41) lim U(z"*, z';1",1') = ~
tll.+tl
fn(Z"*)fn(Z') = e z''~
(2.5.44)
r~0
This shows together with (2.5.36) that for the normal symbol U y we have the initial condition lim UN(z tt*, Zt; Itt,
tl~._~tt
t t) =
1 .
(2.5.45)
50
General Theory
For the harmonic oscillator (2.5.2) we get ]z) = e  i W T / ~ [ze  i w T )
e iTH,r
(2.5.46)
and thus
Uo~c ,tz"* , z"t",t'), = (z"leiT~.dnlz') = eiwT/2(zH]z'e iwT) = ei~T/2exp (z"* e iwT z t) . 2.5.47) Expanding the last result yields
Uosc(z"* ,z ,"t",t') = ~_~ f~(z"*)eiE"T/h f~(z ')
2.5.48)
n=O with the correct spectrum En = hw(n + 89 and a comparison with 2.5.41) gives the expected result osc
Krn n (T) = e 
i
EnT/h
(2.5.49)
5ran 9
To derive the path integral in the coherent state representation, we start from the semigroup property of the timeevolution operator (t" > t > t')
(2.5.50)
u(t", t') = u(t", t) u(t, t') and obtain U(z"*, z ' ; t " , t ' )
=
(z"lU(t",t')lz')
I with c = tk+1  tk := T / ( N + I), T = t"  t' > 0 * 1 = z"* For the timeevolution fixed, to = t~,tN+l = t H, zo = z ~ and Zg+ operator we have for small c U(tk+l, tk) = 11 ~i c H ( a t , _a;tk) + O(e 2) 9
(2.5.52)
Let us assume, without loss of generality, that the Hamiltonian in Fock space is already given in normal ordered form, so that we obtain for the normal symbol of U U N ( z k*+ l , z k ; t k + l , t k )
= 1 ~i e H ( z.k. + l , z k ;
t "k) + O ( e 2)
2.5 Path Integration Over Coherent States = exp
51
[i ~eH(zk+l,zk;t~) , + O(e 2)]
(2.5.53)
and thus for the kernel of U in the space of coherent states [
U(z;+l, zk;tk+l,tk) = e x p
i
z;§
.
 ~eg(zk+~, zk;t~) + O ( e 2)
]
. (2.5.54)
Inserting the last expression into (2.5.51) we obtain with d#(z) = e Iz12 dzdz* 2~ri the lattice definition of the path integral in the coherent state representation
N ~ dzjdz~
U(z"*,z';t",t')=
lim I I N~oo
2~ri
j=l
N
• exp
N
 E k=l
i
g
]
[Zk[' + E z ~ + lzk  hEeH(z;+ l'zk;tk) k=0
k=0
(2.5.55)
Here we have assumed, as usual, that the terms O(e 2) do not contribute in the limit c + 0 and, of course, that the limit N + oo exists. To interpret (2.5.55) as a path integral, we consider independent complex paths z(t) and z*(t) with z(tk) = zk, z*(tk) = Z; and the boundary conditions
zo = z(t') = z' ,
zN+~* = z*(t") = z"*
(2.5.56)
Notice that we do not require at the end points z(t") = z" and z* (t') = z'*. Indeed, only the values z I and z ' * are fixed in the kernel (2.5.55). The first two sums in the exponent of (2.5.55) can be rewritten as N

E
N
ZkZ * k +
k=l
E
N
Z k* Z k _ 1 + ZN+ * 1 ZN = ZN+ * 1 Z N 
k=l
E
k=l
eZ~ Zk
 ~Z k  1
(2.5.57)
which gives in the limit N + oo, e + 0 r
t It
z"*z(t")  Jr, z*(t)i(t)dt and thus we obtain the following
path integral over coherent states
z.(t,,)=z ,,o r
U(z"*,zt;t",t I) =
x exp
[
/
Vz(t)Vz*(t)
z(t,)=z,
z"*z(t")+ ~
(ihz*(t)i(t) H(z*(t),z(t);t))dt
Here we have introduced the "path differentials"
]
(2.5.5s)
52
General Theory N
Vz(t):Dz*(t) " lim V[ [ dzjdz~ N*o~ ~
JC
(2.5.59)
27ri
Here the following remark is in order: instead of (2.5.57), the first two sums in the exponent in (2.5.55) can also be rewritten as
( (~:~) 21 (z"'z(t")+
z'(t')z') + ~l ff' (~'(t)z(t) z'(t)~(t))dt (2.5.60)
which would seem to lead to a symmetrized version of the path integral (2.5.58). (In fact, this symmetrized version of the path integral (2.5.58) is often used in the literature.) However, this symmetrization is in general not correct within the path integral, since the above integrals have to be considered as stochastic integrals! Only if in the above integral ;?* is interpreted as the forward derivative, ~*(t) = lim~,0 (z*(t + r  z*(t))/e, and ~(t) as the backward derivative, i(t) = lirrk.0 (z(t)z(te))/e, do the above expressions lead to the correct path integral, i.e., the correct lattice definition (2.5.55).
2.5.3 The Forced Harmonic Oscillator and the Feynman Propagator. The Hamiltonian of the forced harmonic oscillator reads in normal ordered form as follows
_ _ ~hw 1 11H(~ t , _a;t) = hwata+
J(t)(a + a_)
(2.5.61)
where J(t) is a real cnumber source corresponding to the driving force E(t) = mJ(t). (From the definitions (2.5.3) one sees that the driving term ~/tim/2w g(t)(_a t + _a) equals mg(t)q.) The kernel U of this system has the following path integral representation
U(z"*, z'; t", t') z*(t")=z"" = ei~T/~
j
(
i
~z(t)Z~z* (t) exp z"*z(t") + ~R[~,
)
(2.5.62)
z(t')=z' where we have introduced the classical "action" (c := ~
)
t II
R[z',z] ih ~Jt, [(~ + iwz  i c J ) z * icJzJdt."
(2.5.63)
Since the path integral (2.5.62) is of Gaussian type, it can be exactly solved by the method of stationary phase
2.5 P a t h Integration Over Coherent States
U(z"*,z';t",t')
53
( *"*ZCl(t")+ ~R[z*,zcl] 9 )
= ei~~
(2.5.64)
Here the "classical path" ZCl(t) is the solution of the equation of motion
5R[z*, ZCl] _
i h(kCl + iwzcl  i cJ) = 0
~z*(t)
(2.5.65)
with the initial condition zcl(t') = z'. The solution of this inhomogeneous equation is given by (t' < t < t") ZCl(t) = e i~o(tt') z' + i e
f:
e i~o(t,)
J(s)ds ,
(2.5.66)
and thus we obtain for the expression in (2.5.64)
z"" ~o~(t") + ~n[z',zo~] = ~"'zc,(t") + iv.[,, J(t)zc,(t)dt t II
/ e iw(t''t) J(t)dt Jt ei'(te) J(t)dt c2 [, at dsJ(t)ei~~
= Z"* e  i w ( t "  t ' ) z I 't i c z " *
i
+icz'
at
J(s) . (2.5.67)
The last integral can be rewritten as follows: t"
t
fit' dt~t, dsJ(t)eiW(ts) j(s) = ~ti"dt ~ti"ds J(t)e i~ = 1
ft"
J(s)O(t s)
t"
(2.5.68) Introducing the
Feynman propagator DF(t) := 2~w[ei~~ o(t) +ei~t o(t)] 1 e_ iwltI 2iw
(2.5.69)
kernel of the forced harmonic oscillator (2.5.70) U(z"*,z';t",t')= Uosr lz"* , z';t",t')Z[J]
we obtain the final result for the
where Uosc denotes the kernel (2.5.47) of the harmonic oscillator without driving force (J = 0), and the generatingfunctional Z[J] (Z[O] = 1) is given by
54
General Theory
Z[J] ~ Z(z"*, z';t",#]J) :=exp
[,fd ['" iV2hwJt '
imft" 2ti
(e
ft"
Jr, dtJt,
iw,,,, ,, z,,,
+e
iw(,_t,, z, ) J(t)dt
dsJ(t)DF(ts)J(s)J
] .
(2.5.71)
The Feynman propagator satisfies the inhomogeneous wave equation
+w 2 DF(t  s) = a(t  s)
(2.5.72)
and has the integral representation
DF(t) =/(oo dw' ei,o't 27r w'2 ~ w~ + i e
(2.5.73)
Here a few remarks are in order i) The Feynman propagator DF, (2.5.69) and (2.5.73), is a simple (nonrelativistic) example of the socalled Feynman propagators which play an important r61e in quantum field theory, in particular as building blocks of the Feynman rules, see e.g. [534]. It should not be confused with the Feynman kernel K(z", z'; t", t') which is, however, also called the propagator (or Feynman propagator) by some authors. ii) It is clear from the integral representation (2.5.73) that the specific regularization corresponding to the above "i erule" is of crucial importance for obtaining the correct Feynman propagator. It is a remarkable fact that the coherent state path integral leads automatically to the correct regularization. This is in contrast to the usual path integral which due to its oscillatory character in physical time leads to (2.5.73) without the i eregularization. To cure this disease, one can proceed as follows. The regularization (2.5.73) can be interpreted as making the replacement w 2 + w 2  i e which leads in the usual path integral for the harmonic oscillator to the replacement
exp
( i ~ "~m
w2x2dt )
exp
(_
m ~ t''
e~/~ ,
and thus gives for e > 0 a Gaussian damping. iii) Taking functional derivatives of the path integral (2.5.62), for instance, ~U/3J(t) ]a=0, one obtains the expectation values of powers of a+_at ~ % but these are completely determined by the generating functional Z[J], see (2.5.70) and (2.5.71).
2.6 Fermionic Path Integrals
55
Finally let us also note the Gaussian integration formula for coherent states. Let A be a (d • d)matrix of a nonsingular quadratic form whose Hermitian part is positive, and z and u stand for vectors of complex numbers. Then
dzdz* ~z'Az+u'z+uz" (27ri)d _
eu*Alu 
d~
(2.5.74)
2.6 Fermionic P a t h Integrals 2.6.1 Fermionic Coherent States. Since path integrals exhibit in a particularly clear way the close relationship which exists between classical and quantum mechanics, it would seem a priori that we would encounter some difficulties when extending the treatment to fermions. Fortunately the relevant construction in terms of an anticommuting algebra has been devised by Berezin [7478], Martin [682], and Schwinger [838]. Fermionic path integrals were introduced by Berezin [76], Faddeev [311], Faddeev and Slavnov [313], and other authors. For a review, see Klauder and Skagerstam [602]. In the context of supersymmetric quantum mechanics see also Berezin et al. [7778], DeWitt [236], and Singh and Steiner [855]. To be specific, we restrict ourselves to a Fermi system with a single spin variable. Generalization to many degrees of freedom is straightforward. Consider a spin89 particle which is described by the Pauli matri2 = 11, [c%,Cry] = 2 i Crz and cyclic perces Cr,,Cry,Crz with o~ = Cr~ = Crz mutations, {Cr,,Cry} := Cr*Cru+ CruCr* = 0 and cyclic permutations. With Cr+ := (~, + i ~y)/2 we can define fermionic annihilation and creation operators, respectively, by _a := ~+ , _at := Cr_ , (2.6.1) which obey the fermionic anticommutation relations {_a,_at) = 11 ,
_a2 = (_at)2 = 0 .
(2.6.2)
Notice that we have [~.,_at] = ~z. As in the bosonic case, we shall construct a Hilbert space of "entire functions" (the related mathematical theory of the corresponding functional analysis is also called superanalysis, e.g. [77, 236, 428, 724, 803] and references therein), where, however, the r61e of the complex numbers z and z* (see Sect. 2.5.1) is now played by anticommuting variables, socalled Grassmann variables q and f/satisfying {~,#} = 0 ,
q2 = 02 = 0 .
(2.6.3)
The most general function of the two variables 7/and f/has the polynomial form f(f/, q)  co + Cl~ + e2~ "9FC3~/] (2.6.4)
(ci are complex numbers). We define derivatives 0 and 0 with respect to and f/, respectively, by
56
General Theory O f ( ~ , r/) = c2  c30 ,
(2.6.5)
0 / ( 7 , r/) = cl + e3r/,
i.e., 0 suppresses r/, while c5 suppresses 0 after having brought the relevant variable to the left. The coherent state representation of a fermionic state Ir is defined as the "entire function" r
= co + c 1 0 ,
(2.6.6)
while the complex conjugate is given by r
= e~ + c~r/ .
(2.6.7)
The scalar product in this representation is defined by (r162 = d~co + d*~el
(2.6.8)
if r = do + d10. In order to derive a fermionic path integral, we have to express the scalar product in terms of an integral over Grassmann variables. The Berezin integrals [7478] are introduced as follows
fdOO=fdr/r/=l,
/dr/1
=/dr/1
=0 ,
(2.6.9)
where dr/and d 0 anticommute. This gives for a general function f(0, 7/)
f dr/ f(O, r/) = Of(O, r/) d o 1(0, r/) = cw
r/)
f d o dr~ f ( o , 71) = OOf(O, r/)
(2.6.10) D
With e ~ = 1  Or/we obtain
/ dOdr/eO"r162
/
d O d r / ( 1  Or/)(d; + d'lr/)(co + c10)
[ dOdr/(d*lC~r/O  d;coOr/)
,I = d*lcl + d~co  + r/ll>
(2.6.13/
n=O,1
such that r
is given for a general Fock state 1r r
:= (r/lr
= En=0,1
c,~ln> by
c.(r/I.>
= ~ n=0,1
= Z
Cn/n(~) "
CO [" C10 9
(2.6.14)
n=0,1
With these definitions we find
(atr
:=
(r/l_atlr = ~ c(r/Iatl n) n=O,1
= co(r/l_a310)= co(r/ll) = c0O = Or
(2.6.15)
and similarly (ar
:= (r/I~lr = c, = c,(r/lO> = cl = 0 r
,
(2.6.16)
which shows that the operator at acts in the space of "entire functions" r by multiplication by 0, while a acts by cq (see the close analogy with (2.5.28) and (2.5.29) in the bosonic case). An arbitrary operator A = A(a, at) in Fock space can be written as A =
~
Im>Am, 9
(2.6.17)
0"~Am,~r/''
(2.6.18)
rB,rl:O,l
Its matrix representation is given by A(O, 7 / ' ) : =
= / dOI I dr/r e  '  / / ,7/ t A(0, r/")B(0", r/') . (2.6.20) If an operator A is given in normal ordered form
58
General Theory _A ~
Ck,(at)k(a) 1
(2.6.21)
k,l=O,1
we define its normal symbol by
AN(o, rl') :=
Ck,( O)h (
(2.6.22)
t .
k,l=O,1
It is then easy to see that the relation between the kernel (2.6.18) and its corresponding normal symbol is given by
A(O, 77') = e q"' AN(o, ~') 9
(2.6.23)
Finally let us mention the formula for the Ndimensional fermionic Gaussian integral for N independent Grassmann variables r/,~,0,~, and Grassmann "sources" Jn, Jn f H dOkdr/kexp
 E
OmArnnrln b Z(OnJn [
m,n=l
k=l
n=l N
= detA.exp (,~_
 z  1mnJn ) . (2.6.24) Jm
This should be compared with the corresponding bosonic Gaussian integral (2.5.74). Notice, in particular, the different powers of det A in the bosonic and fermionic formulae, respectively. 2.6.2 T h e F e r m i o n i c P a t h Integral. One observes that the properties of the fermionic coherent states discussed in Sect. 2.6.1 are very similar to the bosonic case discussed in Sect. 2.5.1. Based on this close analogy with the bosonic situation, one easily derives the fermionic path integral for the corresponding timeevolution kernel
(2.6.25)
U(O",r/;t",t') := (r/'lU(t",t')lr/) . The result is the following fermionic path integral [311,855] q(t")=q"
u(o",
=
[
DO(t):D~?(t)
/ *,2
i/,
~(t')=~'
x exp O"y(t")t ~ ,
(iliO(t)il(t)  H(O(t),y(t);t))dt
= Nlirnoof dON, d~N1.., dOt drll
]
2.6 Fermionic Path Integrals
59
N ie _ "tk)) X exp F/N•N  E ( F/k(rlk  ilkl) + '~H(r/k, rlk1, (2.6.26)
k=l
Here F/k and qk denote Grassmann variables satisfying {~/k, rlt} = {F/k, F/t} = {F/k,ql} = 0 for all k,l; F/k = F/(tk), etc., tk = t ' + ek, e = ( t "  t ' ) / N . The boundary conditions are imposed by requiring q(t) to be fixed at t = t', rlo = 71(t') = 7/, and F/(t) to be fixed at t = t", F/N = F/(t'') = F/". Furthermore H(F/, q;t) is obtained from the Hamiltonian H(a t,a;t) given in normal ordered form by replacing the fermion creation and annihilation operators as at _+ F/, a + q. 2.6.3 T h e P a t h Integral for a Spin~ Particle in a T i m e  D e p e n d e n t M a g n e t i c Field C o u p l e d to T w o External S o u r c e s . Let us consider the motion of a spin89 particle in a timedependent magnetic field B(t) along the z axis coupled to two external timedependent Grassmann sources J(t), J(t) characterized by the Hamiltonian (h = 1) 1 H_ =   ~ B ( t ) ~ z  J(t)c,+  ~_ J(t) .
(2.6.27)
With c,+ = _a, ~,_ = at and a~ = [_a,at] one obtains the following normal ordered form of the Hamiltonian (2.6.27) H(a t, a;t) = B(t)(at_a  1 11)  J(t)a  _atJ(t)
(2.6.28)
in terms of the fermionic operators a, at. Inserting the corresponding kernel H(F/, q; t) into the path integral (2.6.26) and carrying out the integrations at every lattice point we obtain [855]
U(F/",~I';t",t')=
lira exp F/N
N+oo N
(1 icBk)7/0
N
+iF/NECJk k1
H
N
k1
(1ieB')+iEeJkH
/=k+l
k=l
(1icBt)r/~ /=1
Nklk1 ](2/:"  E eyk E eJl H (1ieBm) e x p k=2
t=l
re=t+1
) B(t)dt (2.6.29)
just by using the standard Gaussian integration rule (2.6.24) for the Grassmann variables. The continuum limit of (2.6.29) can be easily written clown as
',=exp{O"exp(i/;
60
General Theory
+iO"~dt,l,,oxp(i~ ~ls~,~ls,ds) +i~ d,'l,,exp(i// ~l,s,~ls,ds)~' dr_
d~ J ( , ) ~ , ( t ~}J(~)
_
B(t)d,
@
at
(2.6.30)
Here we introduced the fermionic Feynman propagator in the presence of an external magnetic field B(t)
DF(t,s) := 19(t s)exp ( i f t B ( r ) d r ) .
(2.6.31)
The timeevolution kernel (2.6.30) has the following general decomposition
U(fl", rl';t",t') = Koo + Kn#"rl' + ~7"Klo + Kozrl' ,
(2.6.32)
where the coefficients Kmn are given by
B(t)dt
Koo = exp
exp
(;; 
dsJ(t)DF(t,s)J(s)
dt
)
,
(2.6.33a) K n = exp K 1 0 = i j,tI
K01
 i
B(t) dt Koo ,
(2.6.33b)
, d t J ( t ) e x p (  i ; )B(s)ds
=z i
(z)
at Y(t) exp
(iz;) 
B(s)ds
Koo ,
(2.6.33c) (2.6.33d)
Koo 9
The Km,~ are actually the matrix elements of the timeevolution operator in the twodimensional Fock space spanned by the two vectors [0} and I1)
K.~.=(mTexp(iftT"H(a_t,a_;t)dt )
n) .
(2.6.34,
Notice that the generating functional (=vacuum persistence amplitude) for the pure fermionic system is given by
Zg[J, J]  Ko0lJ=J=0
C;
 at, dt
, as ](t)Df(t, s)J(s)
)
.
(2.6.35)
2.6 Fermionic Path Integrals
61
2.6.4 S u p e r s y m m e t r i c Q u a n t u m Mechanics. In supersymmetric (SUSY) quantum mechanics [379, 724, 934] one considers the Hamiltonian (units h = 2m = 1 are used in this section) (2.6.36) which corresponds to the "quantum Lagrangian" s
= ~ 1q .2  ~ [1V ( q ) ] 2+iat_s
1[at, _ a]V'(q) _ ,
(2.6.37)
where a and _at are fermionic operators. The potential V(q) is called the superpotential. For the choice V(q) = q the above system describes two noninteracting bosonic and fermionic oscillators. Notice that the fermionic part in the Hamiltonian (2.6.36) can be identified with the Hamiltonian (2.6.27) of a spin89 particle in a timedependent magnetic field B(t) = V'(q(t)) without external sources. If one wants to calculate the path integral for SUSY quantum mechanics, one first has to compute the fermionic path integral for the pathdependent "magnetic field" V'(q(t)) and then in the second step to integrate the result over the bosonic degrees of freedom. In such calculations, the commutator in the Lagrangian (2.6.37)is usually replaced by 89 a] + r 1 6 2treating r r as "classical" Grassmann variables, and the fermionic path integral is set equal to the fermion determinant (see (2.6.24))
f:D(b(t):Dr162162
] _= det ( i
d _ V'(q(t)))
(2.6.38)
However, the correct replacement [855] of the commutator to be used in the 1 In the literature, (2.6.38) is fermionic path integral is 89 , _a] + r 1 6 2 3" essentially used as the defining equation of the fermion determinant which is later evaluated not from the path integral but by solving an eigenvalue problem with appropriate boundary conditions. Since the determinant is finally normalized by hand, one arrives at the correct kernel even though the correct quantum Lagrangian has not been used. Actually the path integral (2.6.38) does not stand well defined without specifying the boundary conditions. Depending on the applications one is interested in, one obtains different expressions for the fermion determinant which, however, can be exactly obtained if one uses the correct path integral (2.6.30). If one is interested in the trace of the timeevolution operator [205, 391], then we easily obtain [855] from the path integral (2.6.30) putting J = ] = 0 and B(t) = V'(q(t))
62
General Theory d
_
= [Koo +
f tll
V'(q(t)))trace:=Tr[Texp(i]t,
dee ( i ~~
dt _HF) ]
Kll]:=j=o
U(fl, o;t",t')b=:=o
= / dr/dr/e  ~
at y ' ( q ( t
= 2cos
,
(2.6.39)
where H F is the fermionic part of the Hamiltonian (2.6.36). Notice that the argument  0 of U in (2.6.39) is the origin of taking antiperiodic boundary conditions in the evaluation of the fermionic determinant in the earlier works. The Euclidean version of (2.6.39) is identical to the result obtained in [205, 3911 9 The trace of the pure fermionic system in the case of a constant magnetic field B(t) = 2w > 0 can be easily obtained from (2.6.39) Tr T e x p
 i
= 2 cos(wT) = e i~T + e i~T
(2.6.40)
leading to the identification of two energy levels with energy E0 =  w and E1 = w as is expected for a spin89 particle in the presence of a constant magnetic field. If the path integral (2.6.38) is used to define the generating functional as is the case with the Nicolai map [724, 855], it must be interpreted as the vacuumvacuum transition amplitude, and thus we obtain det
i ~d
_
V'(q(t)))
/vacuum "
Koob_j_o
=  [ d O d y e  ~ OU(O, TI;t",t')tj=j=oq
This result for the fermion determinant turns out to be exactly the inverse of the bosonic Jacobian for the Nicolai map as evaluated by Ezawa and Klauder [308, 309] using the Stratonovich prescription. One therefore obtains exactly the cancelation of the fermion determinant and the bosonic Jacobian under the Nicolai map. Finally, let us discuss the Witten index [935] A := Tr(I1)E, where (  ll)g = 112_at_a is the fermionnumber operator, and which has been introduced as a measure of supersymmetry breaking. The regularized version is defined as follows (/~ > 0)
:_
[(
Y]
2.7 The Path Integral in Spherical Coordinates
63
q(#)=~ = /_~d x
7~q(s)exp (  ~Xj~o~[02+
/
[( )FeaU ],
V2(q(s))]ds}
q(O)=z
(2.6.42) where the trace over the bosonic degrees of freedom has been converted to the path integral form. The remaining trace over the fermionic degrees of freedom is immediately obtained from (2.6.33) as Tr [(II)Fe~H~ ] =det (d~ss +
V,(q(s)))S_trace
=
f  / d ~ d ~ ? e 07 U(~, T];~
i~
O~J~J~
0
= 2sinh [1~o~ ~ dsV'(q(s)) ]
(2.6.43)
Here the superscript E stands for Euclidean form and Strace for supertrace. It is to be noted that in this case we have kept the arguments of U unchanged which implements periodic boundary conditions in calculating the determinant. Inserting (2.6.43) into (2.6.42) the complete expression for the Witten index is obtained. 2.7 The P a t h Integral in Spherical Coordinates We consider the DdimensionM path integral (2.2.20) (x E IRD)
x(t')=x' = lim
YI
dxk
k1
i m / xj2 + x ~j _ ,  2xj' x HN e x p J[2~t j=l
xj_I)~VCxj)] ie
I
J
(2.7.1) (Notice that in the second product j has been replaced by j  1, but that instead of V(xj_l) we still write V(xj) which makes, however, no difference under the path integral; indeed one could also use a more symmetric formulation where V(xj) is replaced by (V(xj) + V(Xj_l))/2.) Let Y(x) be a function of JxI only, Y(x) = V(Jxl), and introduce Ddimensional spherical coordinates
64
General Theory X 1 =
rcos01
z 2 = r sin 01 cos 02 z a = r sin 01 sin 02 cos 03 z D  1 ~~ r sin 01 sin 0 2 . . . sin
(2.7.2)
0 D 2 cos ~o
Z D = r sin 01 sin 0 2 . . . sin 0 D  2 sin ~ ,
where 0 _< 0 v _< r (v =
(~'~D=l(zv)2)l/2>_ cos r162
1,...,D
2), 0 _< ~  0 n  1
V(r).
0, thus V ( x ) =
__~ 27r, 7" ~
We use the addition theorem
= cos 01 cos
D2
+ ~
c o s 0 ? +1 cos0~n+l
m=l
fi
D1
s i n 0 ~ s i n 0 ~ + H sin0~sin0~ ,
n=l
(2.7.3)
n=l
where r is the angle between two Ddimensional vectors x l and x2 so that Xl .x2 = rlr2 cos r The metric tensor in spherical coordinates is
(gab) =
diag(1, r 2, r 2 sin 2 0 1 , . . . , r 2 sin 2 0 1 . . . sin 2 0 D2) .
(2.7.4)
If D = 3, (2.7.3) reduces to: cos r
= cos 01 cos 02 + sin 01 sin 02 cos(~l  ~2) 9
(2.7.5)
The Ddimensional measure dx expressed in spherical coordinates reads D1
dx = rDldrd~2 ,
d• = H
(sinOk)Dlkdok "
(2.7.6)
k=l
d~2 is the ( D  1)dimensional surface element on the unit sphere S D  1 and 12(D) = 2rD/2/F(D/2) is the volume of the unit S ~ The determinant of the metric tensor is given by g :=
det(gab)
=
D1 r D  1 H (sin o k ) D  l  k k=l
)2 .
(2.7.7)
The path integral (2.7.1) can be rewritten in spherical coordinates as follows K ( x " , x ' ; T ) = K ( r " , 1 2 " , r', f~';T)
limoo (\ ~1~ m ~ND/2Nlfo~176 = N~ r 191 k drk f d~2k ] H k=l
N Jim t 2 x H exp + j=l L
2 j_x 2 jr _l cos

(2.7.8)
2.7 The Path Integral in Spherical Coordinates
65
For an explicit evaluation of the angular integrations we need the expansion formula for plane waves [413] e r162176162 =
F(u)E(l+u)I,+u(z)Cr(cosr
,
(2.7.9)
I=0
(valid for any v r 0 ,  1 ,  2 , . . . ) , where C~ are Gegenbauer polynomials and I~ modified Bessel functions. The addition theorem for the M linearly independent real surface (or hyperspherical) harmonics S~ of degree 1 on the s o  l  s p h e r e has the form [303, Vol.II, Chap. IX]: M
1 21+D2 D2 J2(D) D   2 C ' 2 (cos r
Es~(allS~(il2)
(2.7.~o)
,
/~1
M = (2l + D  2)(l + D  3)!/l!(D  3)!, with the unit vector il = x / r in 1RD. The orthonormality relation reads (2.7.11)
dil S~ (il)Sl~, ' (ll) = 5u'5~u' 9 Combining (2.7.9) and (2.7.10) we get the expansion formula D._.~2
M
e z("'"=) = eZr176162 = 2 .
S~'(~I)S~'(a2)I,+ D;2 (z) 1=0 p = l
(2.7.12) The angular integrations can then be carried out and the path integral (2.7.8) in spherical coordinates becomes K (r", gl", r I, ill; T) = ( r'r')~a E
E
Sr162
1=0 p.=l
• r I exp [ ~ ( r j
lim N.oo
m ~eh
rkdrk k1
+ r~_l) ~V(r~) ~,+~r~ ~ r ~ r ~ _ l
.
j=l
(2.7.13) Therefore we can separate the radial part of the path integral (partial wave expansion) K(r",12",r',~';T)=
( r t r " ) L ~  ~ 21 + D  2 D2 ~(D) l=o D  2 Cg(c~162
(2.7.14) with the angularmomentum dependent radial kernel Kl given by the radial path integral [865]
66
General Theory
Kl(r",r';T)= lim x
=
pl+_~_.x[r2] 9exp
( ,,, t, 7777 )
[5c hj=i
/a/dudvexp [i(qq)"u +i(pp__)"v]Iqj_ i ) 1 f~ [i ie ] (2rrh) D D dpd exp ~ p j  Aqd  ~Hefr(pj,clj) ,
(2.8.24)
where ftj = 89 + qd:) is the jth midpoint coordinate and Aqd denotes the difference Aqj = qj  q j  i . The effective Hamiltonian is given by 1
ab
(2.8.25)
Heff(pd, (lj) = ~~mg (Clj)Pa,jPb,j + V(ctj) + AVWeyl((tj) 9
Inserting (2.8.24) in the composition law, see (2.2.2), we obtain the Harailtonian path integral [135, 221, 331, 367, 563, 698, 786] in the midpoint prescription 4
N1
N
k=l
l=l
K(q",q';T)= N~oo lira II /a ~ dqk'lI/a ~ (2~h) dp~D x exp
q(t")q"d f D(q(t), p(t)) exp q(tt)=q,
{ ,t''hi ~t
[P" 6  Herr(p,q)]dt}
4
(2.8.26)
Note the asymmetry in the above integrations over q and p: while there are only N  1 integrations over q, there are N integrations over p; furthermore, the qpaths are fixed at the end points, as usual; the ppaths, however, are not restricted at the end points. Integrating out the momenta pj by means of the Ddimensional Gaussian integral (2.3.5) yields the Lagrangian path inte9ral in the midpoint prescription ( M P = MidPoint): K ( q ' , q';T)
t"
q(t")=q"
q(t')=q' l=l
k=l
4 Note the slight discrepancy in the jsummation in comparison to (2.2.6). The reordering of the summation affects, however, only terms of O(e2) which can be neglected.
72
General Theory
(2.8.27) with the effective Lagrangian /:eft(q, Cl) = 2gab(q)qaq b  Y(q)  AYweyl(q) 9
(2.8.28)
The midpoint prescription arises here in a very natural way as a consequence of the Weyl ordering prescription. It is a general feature that ordering prescriptions lead to specific lattices and that different lattices give different quantum corrections A V oc h 2. For a thorough discussion of further subtleties of the path integral, see e.g. Babbit [43], Keller and McLaughlin [571], and Nelson [721]. To prove that the path integral (2.8.27) is indeed the correct one, one has to show that with the corresponding shorttime kernel
(
m
K(q/,q/_1;e) = \ ~ ]
~DI2
[g(qJ')g(qJ)]l/4~/g(~lJ )
the SchrSdinger equation (2.8.1) follows, e.g. [464, 736], from the timeevolution equation ~'(q",t") = / R o dq' ~
g (q", q' ; t", t')C~(q', t') .
(2.8.30)
Let us emphasize that this procedure is nothing but a formal proof of the path integral. A rigorous proof must include at least two more ingredients i) One must show that in the limit N + oo the path integral representation for K is in fact the matrix element of the timeevolution operator U(t", t') for all ~ E 7/ (7i: relevant Hilbert space). ii) One must show that the domain 7) of the infinitesimal generator of the kernel K is in fact identical with the domain of the Hamiltonian corresponding to the SchrSdinger equation (2.8.1), i.e., the infinitesimal generator is the (selfadjoint) Hamiltonian. Concerning our formulation of the path integral for the quantum motion on curved manifolds, one might ask the following questions: i) Is it really necessary to include the quantum potential A V ? In our approach the emergence of the quantum potential AV is absolutely unavoidable. Once the quantum Hamiltonian is defined to be given by the LaplaceBeltrami operator, a particular quantum potential emerges as an unavoidable consequence of the chosen ordering prescription of position and momentum operators, and enters the evaluation of the shorttime
2.8 The Path Integral in General Coordinates
73
matrix .element of the timeevolution operator. However, one is still free to make further manipulations in the lattice formulation of the path integral in order to cast the effective Lagrangian into a convenient form, and there are known examples, indeed, where the quantum potential is cancelled by another contribution. In fact, a path integral formulation without an h 2 quantum potential can be developed. According to Kleinert [611,613] it is based on the equation for the straightest lines, i.e., F ~ 5~
0
(2.8.31)
on some Riemannian manifold IM with coordinates q. The (full) affine 9 )~ 9 ,k ~,~ )~ )~ connectmn F~, is constructed by F~, = Fs + K , , , where F~, are the usual Christoffel symbols, and K,~x := S,~x  S , x , + Sx,, is the contortion tensor with S,,x the torsion tensor. The result is finally the lattice formulation of the path integral as in Sect. 6.1.1.5. Therefore this path integral formulation includes torsion in space, respectively spacetime. ii) Is it always necessary to use the lattice definition of path integrals ? Our discussion is based on the timesliced definition of the path integral, as it has been introduced by Feynman and is usually used in the physics literature9 It is true that all basic path integrals have been obtained from a timesliced evaluation, and eventually taking the proper limit N + co. The emergence of potentials o( h 2 in the evaluation of the matrix elements of the shorttime propagator is closely related to the stochastic nature of the Feynman paths. Of course, for building a conceptual general calculus, one may ask whether it is possible to define path integration without a limiting procedure. According to DeWittMorette (cf. the above cited literature and other authors) this is possible. In fact, as long as Gaussian path integrals are concerned, things work out perfectly well, and this theory allows a comprehensive formulation of the spacetime transformation technique of Sect. 2.10.3, cf. [152, 343, 344, 943]. Recently, Cartier and DeWittMorette have proposed a method [147, 240] to avoid the timeslicing definition for path integrals on curved manifolds. The method has been used by LaChapelle [628] to find a spherical path integral formulation. It is obvious that the mathematical questions coming along with the path integral have attracted the attention of many authors. Among them are A1beverio [1116, 1820], Arthurs [34], Sabbit [43], Cohen [198], DeWitt [235], DeWittMorette et al. [147, 237248,629,710712], Dowker and Mayas [265, 266, 690], Elworthy and Truman [299,300], Fischer et al. [343345], Garrod [367], Gervais and Jevicki [389], Grosche and Steiner [464], Kac [554], T.D.Lee [645], McLaughlin and Schulman [665], Marinov [676], Mizrahi [698], Omote [736], Papanicolaou [758], Prokhorov [787], Steiner [863,865], Truman [897], and many others who can be found in the literature.
74
General Theory
Furthermore, some instructive proofs can be found in the books of Simon [854], and Reed and Simon [794]. Also due to Albeverio et al. [1116, 1820] is a wide range of discussions to formulate the Feynman path integral without the delay of going back to the definition of Wiener integrals. Another approach is due to D e W i t t  M o r e t t e ("Definition Without Limiting Procedure") [237, 246, 699, 701, 711, 712]. Here this formalism is used to set up a rigorous formulation of the semiclassical expansion [238,699705]. 2.8.3 P r o d u c t O r d e r i n g . To develop another useful lattice formulation for path integrals, we consider again the generic case [422]. We assume that the metric tensor gab is real and symmetric and has rank (gab) = D, i.e., we have no constraints on the coordinates. Thus one can always find a linear transformation C : q~ = Cabyb such that the kinetic energy term in the classical Lagrangian s is equal to (m/2)Aabijaid b with Aab = ~ c Ctacgcdcdb and where A is diagonal. C has the form Cab = u (b) with u(b) (b E {1, . . . , D}) being the eigenvectors of g~b, and Aab = ~]~cf~Sac6br where f~ r 0 (a e { 1 , . . . , D}) are the eigenvalues of g,~b. Without loss of generality we assume f~ > 0 for all a E {1, . . . , D}. s Thus one can always find a representation for g,,5 which reads D
gab(q) =
h~(q)hb~(q)
E
(2.8.32)
9
Here the hab = Cool, Cob = u a)Lu! b) are real symmetric D • D matrices and satisfy hab hbc : 6ac. Because there exists the orthogonal transformation C, (2.8.32) yields for the ycoordinate system (denoted by IMy):
Aab(Y) = E f2e (y)6acSbc 9
(2.8.33)
c
(2.8.33) includes the special case gab = Aab. The LaplaceBeltrami operator expressed in terms of the inverse matrix h ab reads on IMq, i.e., in the original coordinates q (h = det(h~b)),
abe
OqaOq'~+
Oqa
+
Oqa
~qb
'
(2.8.34) and on IMu
1 r~bb0 2
&b :a a) 0 ( :ba  2ff
+\h
]
(2.8.35)
s Actually, the case f~ < 0 is possible. This can be seen if one considers pseudoEuclidean (Minkowski) spaces with indefmite metric. All the relevant formulae are also valid in this case, see e.g. [447].
2.8 The Path Integral in General Coordinates
75
With the help of the momentum operators (2.8.3) we rewrite the Hamiltonian in the "product ordering" form ( P F = Product Form)
ha~(q)Papbhb~(q) + V(q) + AVpF(q) ,
H = 2ram h2 A~'LB+ V(q) = ~ml Z abe
(2.8.36) with the well defined quantum potential
h~ ~ [ AVpF(q) = ~m
[4haChbC ,ab + 2haChbc h,ab h
bc
hb
ac
be
h ah b
On IMu the corresponding A V ~ ~ is given by
~.(y)= h' ~
AVpF
7. \
1 [(fb,a~2_4f.,aaSbb
fa D
/b ] + 2 k 7b ],oJ "
(2.8.38)
The expressions (2.8.37) and (2.8.38) look somewhat circumstantial, so we display a special case and the connection to the quantum potential AVweyl. i) Let us assume that Aab is proportional to the unit tensor, i.e., A~b = f23~b. Then AVp~F~ simplifies to 2D  2
AVOn'(y) = h ~
W
~
(4 
D)f2,~ + 2f. f, aa 74
(2.8.39)
a
This implies that if the dimension of the space is D = 2, then the quantum potential AYpF vanishes. ii) A comparison between (2.8.37) and (2.8.19) gives the connection with the quantum potential corresponding to the Weyl ordering prescription: h2 A VpF(q) : AYWeyl(q) Jr ~mmZ
~/2haChbC,"b
hac,.''l'bc,b hac,bhb~,a:] 

abe
(2.8.40) In the case of (2.8.33) this yields:
AVp~F~ = A Vw;y + ~h~ Z ] 2 _..f~L.f,,,,
(2.8.41)
These equations often simplify practical applications. More special cases have been listed in [447].
76
General Theory
Evaluating the shorttime matrix elements then gives the
Lagrangian path
integral in the "productform'definition K (q", q'; T)
q(t')=q' :=
lim
m
dqk
(~ehac(qj)hbc(qj_l)Aq~Aqb e V ( q j )  cAVpF(qj))
x exp [ h  ~
j=l (2.8.42) In particular in Chap. 6, we use throughout the path integral formulation of (2.8.42) if not otherwise and explicitly noted. 2.9 T r a n s f o r m a t i o n Techniques 2.9.1 G e n e r a l R e m a r k s . Let us consider a onedimensional path integral
=
where AV =
J
~(t,)=x'
,
,
(2.9.1)
h2(F2 + 2F')/8m denotes a quantum potential due to a non
trivial metric ~ = e l i r(~')d~', x is a real variable with range  c ~ < a < x < b < c~. It is now assumed that the potential V + AV is so complicated that a direct evaluation of the path integral is not possible. We want to describe a method for transforming a path integral to calculate K or G, respectively. This method is called the spacetime transformation technique and was originally developed by Duru and Kleinert [279,280] in order to treat the path integral for the Coulomb potential (based on a time transformation, see e.g. [878, p.201], and the KustaanheimoStiefel transformation, both well known in astronomy [627, 464470, 507, 630]). However, this was done in a more or less formal manner, and it did not take long before the technique was refined by Inomata [514,516], Duru and Kleinert [280], and Steiner [863865]. This was followed by a huge amount of path integral treatments and discussions, see Anderson and Anderson [26], Bernido, CarpioBernido and Inomata [85, 142, 143], Cai and Inomata [127], Castrigiano and Sts [152], Chetouani et al. [177, 179, 187189], Fischer, Leschke and Miiller [343,344], Grinberg, Marafion and Vucetich [417,418], Refs. [434,436,437,464470], no
2.9 Transformation Techniques
77
and Inomata [495], Inomata [5161, Inomata and Kayed [525] (Dirac Coulomb problem), Junker [549], Kleinert [608], Kubo [623], Lawande and Bhagwat [642], Pak and Sbkmen [743,859,860], and Young and DeWittMorette [943]. In order to understand the basic features, let us start by considering a kind of Legendre transformation of the general onedimensional Hamiltonian:
g ~ :  2 m
~
+ r(~)
+ v(~)  E
(2.9.2)
which is a Hermitian operator with respect to the scalar product (fl, f2) = f a x v ' ~ f ~ ( z ) f 2 ( z ) . Introducing the momentum operator h(d 89 ) P~ = "~ ~z + F(z) ,
F(z) 
dln ~  ) dz
'
(2.9.3)
E
(2.9.4)
_HE can be rewritten as
+
+ ~m[F~(z) + 2F'(z)]

with the corresponding path integral ("promotor" [517,528]) KE (x", z'; T) = xt;T), where K denotes the path integral (2.9.1). Let us consider the transformation z = F(q), and let G(q) = F[F(q)]. Then we get for the operator _HE expressed in the variable q, which we denote by HE,
e iTE/h K ( z " ,
HE 
2m F'2(q) ~q~+ G(q)F'(q)
F'(q) ] dqJ +V[F(q)]E . (2.9.5)
With the constraint f[F(q)] = F'2(q) we get for the new Hamiltonian I~I := f~E: 1
2
~I = ~mp q + f[F(q)l[V(F(q))  E] + AV(q) ,
(2.9.6)
where /~(q) = G(q)F'(q) f"(q)/F'(q), pq = ln(d/dq+/~/2), and AV(q) denotes the well defined quantum potential
h2[(F''(q)~22F''(q) ] F'(q) + (G(q)F'(q))2 + 2G'(q)F'(q)
AV(q)=~mm 3 \ F'(q) ]
(2.9.7) The path integral corresponding to the Hamiltonian I~I is q(s ")=q"
~'(q",q';s") =
/
Vq(s)
q(O)=q'
x exp ~
~q I[F(q)][V(F(q)) E ]  AV(q) as
. (2.9.8)
78
General Theory
Here the "pseudo time" variable s = s(t) is defined by s(t) = f~ dr/F'~(q(r)), with s(t I) = O, s(t")  s" (see below in Sect. 2.9.3). Note that for G = O, A V is proportional to the Schwarz derivative of the transformation F. As is easily checked, we can derive from the shorttime kernel of (2.9.8) in the new "time" s via the timeevolution equation
~(q', s') = / K(q", q'; s", s')~(q', s') dq'
(2.0.0)
the timedependent Schrhdinger equation
II ~P(q, s) = i li~sgr(q, s)
.
(2.9.10)
2.9.2 P o i n t  C a n o n i c a l T r a n s f o r m a t i o n s . The crucial point is now the lattice derivation of ~((s '1) and the relation between K(s") and K. Let us consider the path integral K in its lattice definition (we assume J  1) K ( x " , x ' ; T ) = N.+oolim( 2 ~ ) x 1I j d x k k=l
ND[2 (2.9.11)
exp j1
To transform the coordinates x into the coordinates q by means of the point canonical transformation [180, 260, 314, 359, 389, 464, 512, 736, 786, 848] x = F(q) with Fa(q) denoting the components of F, we use the socalled midpoint expansion method. We must expand any dynamical quantity in question which is defined on the points qj and qj1 of the jthinterval in the lattice version about the midpoints (~j = 89 + qj1) keeping terms up to order (q~  q~_l) 3. Furthermore one must use the path integral equivalence relations [30,235, 389] (here we incorporate a metric gab with its inverse gab in order to state the general formulae):
AqaAq b * i ehg ab m
AqaAqbAqOAqd" i
(2.9.12)
[gabgcd+gaog d +ga g c]
(2.9.13)
Aq a Aq b AqCAqd Aq e Aq] 9 _~
 g ab g ce gdf _[_gab gCf gde gab gcd gcd _]_ffac.qbd gel b gad gbc gef tk gCd gae gbf _{_gcd gaf gbe b gaC gbe gd! + gaC gbf gde _{_gbd gae gCf 4gbd gay gCe + gad gbe gCf + g ad g b] g c e t g bc g ae gdf .~ gbc.qaf gale1 (2.9.14)
2.9 Transformation Techniques
79
These relations are sufficient for all practical purposes. We now have:
AFa(qj)  Fa(qj)  Fa(qj_l) = Fa((lj + ~~J )  Fa((tj
~~J)
: ~qJ~mOFa(q)~qm q:% + 1AqrAq~Aq~ COqmO~(~)lcOqk oqn q:% + . . . .
(2.9.15) Here Aq~ denotes integral
q~  qjm_1. This
gives the coordinatetransformed path
g ( F ( q " ) , F(q'); T) = [F;q(q')F;q(q")] 1/2
( m ~ND/2NI/ N N.}r ~ k ~ J H god"H F;q(qk) 1=1 k=l
x lim
~eAqj Aqj Fm(qj)Fn(qj )  eV((lj)
• exp j=l
d~2 [ a 
al(F;q,m(qj)F;q,n((lJ) \ F;~(~tj)
8m (Fm(qj)F,~(qj))
F;q,mn(~lJ ) ) F;q(~lj)
 [F;q(q')F;q(q")]l/2 q(t")=q"
•
/ i
t"
f DMPq(t)F;q(q't)exp (h ~' ]kfm'~a y r m r ,~a n q"nq"m  V(q) q(tl)=qI 8m (F~nFa)i
F;q
q,n
F;q,mn a Fa,nkl F_i F;q _ F ,m mnkl
"
(2.9.16) Here F ~ = OFa/Oq m, F;q = IF;ql the Jacobian, and F~,kl 1 is defined by (F?m = F~(~lj) etc.)
a Fj,n/ a xl[l~a Fr~nkl(qj) = ~(Fj,m ~j,k FjaJ,"~I I

+ (F~,,,~F~,k)I(F~,,F~,n)1 + (F~,mF~,t)I(F~,kF~,,) 1 .
(2.9.17)
The path integral (2.9.16) has the canonical form, i.e., the usual coordinate transformation from flat space to a nonlinear coordinate system gives the quantum potential AVweyl without the curvature term. In particular (2.9.16) takes, in the onedimensional case, the form (F'  dF/dq) [223, 314, 359, 380, 389,392,464, 512,593, 786,848]:
80
General Theory
K(F(q'I),F(ql);T) = [Fl(ql)F'(ql')]l/2
lim
H dqt ( m N/ NIf 1=1
N
di 2 ~ m n ~eAqj Aqjn F 12 (qj)  eV(qj)
• H F'(qk) exp k=l
(qj)
8m F'4(qj)
j=l
(2.9.18) It is not difficult to incorporate the explicitly timedependent onedimensional coordinate transformation x = F(q, t) [180, 440, 469, 773,872]:
K(F(q", t"), F(q', t'); t", t I) = [Fl(q",t")F'(q',t')]1/2A(qll, ql;t",t')ff[(q",ql;t",tl) , with the prefactor
(F'(q, t) = OF(q, t)/Oq, F(q, t) = OF(q, t)/cOt, etc.)
A(q", ql; t", t I)
,,
/
q F'(z,t")F(z,t")dz 
= exp
(2.9.19)
F'(z,t')F(z,t')dz
)
,
(2.9.20)
and the path integral/~(t", t I) is given by q(t")=q"
_fs
= /
VMpq(t)F'(q,t)
q(t')=q'
xexP{h~ti"[~(F'2(q,t)(12+k2(q,t))V(F(q,t)) 8m F'4(q,t)
m / (Fl(z,t)F(z,t) + (2.9.21)
For the case F'(q, t) # 0 this can be simplified to
[f (F(q", t"), F(q', t'); t", t') q(t~
/
:DMpq(t)F'(q,t)exP[h~ t " ([__~F,2(q,t)(12 V(F(q,t,)
q(t')=q'
8m F'4(q,t)
m
(2.9.22)
2.9 Transformation Techniques
81
In the Ddimensional case x = F ( q , t ) the corresponding transformation formulae are considerably more complicated and are not stated here, cf. [774]. 2.9.3 T h e M e t h o d o f S p a c e  T i m e T r a n s f o r m a t i o n s . It is obvious that the path integral representation (2.9.21) is not completely satisfactory. Whereas the transformed potential V(F(q)) may have a convenient form when expressed in the new coordinate q, the kinetic term (ra/2)F~2q 2 is in general nasty. Here the socalled "time transformation" comes into play which leads in combination with the "space transformation" already carried out to general "spacetime transformations" in path integrals. The time transformation is implemented [279,280,464,495,514, 516,613,773,863,865,870] by introducing a new "pseudotime" s. To do this, one makes use of the operator identity
1 1 _H  E  L (~, t) ~ (~, t)(_H  E ) A (~, t) fz (x, t ) ,
(2.9.23)
where _IzIis the Hamiltonian corresponding to a path integral K, and ft,, (x, t) are functions in x and t, multiplying from the left or from the right, respectively, onto the operator (_H  E). For the new pseudotime s one assumes that the constraint $ tt
j o d s f t (F(q(s), s))fr (F(q(s), s)) = T = t"  t I
(2.9.24)
has for all admissible paths a unique solution s" > 0 given by t"
tH
ft(x, r)fr(x, r) =
F'2(q(r), r)
Here one has made the choice It (P(q(s), s)) = L (P(q(s), s)) = F'(q(s), s)) in order that in the final result the metric coefficient in the kinetic energy term is equal to one. A convenient way to derive the corresponding transformation formulae is to use the energydependent Green function G of the kernel K. Let us first consider the timeindependent case. For the path integral one obtains by simultaneously implementing the point canonical transformation and the time transformation the following transformation formulae K(x", x'; T) = ~ i r
2rcidEe
C(q",q';E) = ~ [ F ' ( r 1 6 2
i ET[?i
G(q", q'; E) ,
lt/~
r ~176
J0 d s " K ( r 1 6 2
with the transformed path integral/s given by
(2.9.26) s")
,
(2.9.27)
82
General Theory
/~(q", q'; s") = lim
i
\ 1/2 NI
N+cr
.
dqk
xexP[hj=~(~e(Aqj)2~F'~(~tJ)(V(F(~tJ))E)eAV(~tJ))] q(s")=q"
s"
/7)q(s)exP[hfo
(202F~2(q)(V(F(q)E)AV(q))ds]
,
q(0)=q'
(2.9.28) with the quantum potential AV given by h 2 [ F ''2 F"'~ AV(q):~mm~3~2~) .
(2.9.29)
Note that AV has the form of a Schwarz derivative of F. For the timedependent case the formulae must be modified slightly, and we obtain the spacetime transformation formulae r
"I 1/2
K(x",x';t",t') = [F'(q",t")F'(q',t')J x
/dE


2zri e

A(q",q';t",t') ,
i ET/h G(q", q'; E)
(2.9.30)
G(q", q'; E) = ~i fo ~ ~[(q", q'; s")ds" ,
(2.9.31)
with the path integral K(s") given by q(s")=q"
h'(q", q'; s") =
/
Vq(s)
q(0)=q'
~q  F'2(q,s) W(F(q,s)) E  AV(q,s)
•
ds (2.9.32)
AV denotes the quantum potential
3y(q,s)
=
~ ,, h2 ~3 F'2(q's) 2 F lllltq, 8 }~,,,2~
[q
r
!
(2.9.33) The rigorous lattice derivation is far from being trivial and has been discussed by several authors. Recent attempts to put it on a sound footing can be found in Gastrigiano and St/irk [152], Fischer et al. [343, 344] and
2.9 Transformation Techniques
83
Young and DeWittMorette [943]. In terms of stochastic processes the timetransformation is formulated as follows:
=
~e~/h
ds"
a0
[
\
(n%q,)
VW[q]~(FCq(s))x')
]
(q)(u(F(q)  E) +
(2.9.34)
Here C(IR, x') denotes the set of paths in N which start at x ~ at t', the 5functions describe the boundary condition, and •W[x] is the stochastic measure for the Feynman process in real time, or the Wiener process in imaginary time after a Wick rotation. Finally, let us consider a pure time transformation in a path integral. Set G ( q ' , q'; E) = k / f ( q ' ) f ( q " )
• gi~0~ as" / q" exp (  is"V/j(U  E ) V ~ / h ) q' } ,
(2.9.3S)
which corresponds to the introduction of the new "pseudotime" s" = f~"ds /f(q(s)), and assume that the Hamiltonian H_ has the product ordered form. Then ~
G(q",q';E) = ~,i(f' f"~89
(2.9.36)
with the transformed path integral
q(s")=q" /~'(q", q'; s") =
f
:Dq(s)V/~
q(O)q j
"~hachcbq q
AVpF(q)
E
ds
.
(2.9.37) Here hac = hac/v/f, ~ = det(h~r and (2.9.37) is of the canonical product form. It is obvious that only in certain particular cases can the metric terms be transformed to unity. The transformation to unity is always possible in one dimension, though, and this is actually the case in most papers found in the literature. 6 Note the parametrization invariance in the case D = 2, which can be interpreted as a gauge transformation, cf. [356].
84
General Theory
2.9.4 SpaceTime Transformations in Radial Path Integrals. In Sect. 2.7 we discussed the path integral in spherical coordinates for systems that are invariant under rotations, i.e., for which the Lagrangian reads (limr.o[r2V(r)] = O) i'n. 2
s = ~x  V(r)
(2.9.38)
with r  Ixl. Restricting ourselves, without loss of generality, to the threedimensional case, D  3, we have for the Feynman kernel the following "partial wave" expansion (see (2.7.14)) O0
K(r",~2",r',12';T) 
1
4.r,rS#
~~(21 + 1)K,(r",r';T)Pl(cosO)
(2.9.39)
/0
where ~9 denotes the angle between ~ ' and 12" and the radial kernel IQ is given by the radial path integral (see (2.7.15))
Kl(r", r'; TIV ) =
0, if one replaces the phase factor  i(lr/2)(n + 89 by  i(~r/4)  i(~r/2)u(T), where u(T) := [wT/~r] is the socalled Morse index (see Chap. 5, [x] denotes the integer part of x). One then obtains the correct relation lim
T~(nTr/w)+
Kosc(Z", x'; T) = einTr/2 ~(X,t
_
(1)'x')
(3.2.23)
which generalizes the initial condition 2.1.23) for T + 0 +. Finally we observe that the path integral for the Ddimensional harmonic oscillator can be expressed in terms of the MoretteVan Hove determinant (see Sect. 5.2, remark v)) as follows
1 I
Ko~(x",x';T)  (2rrih)D/2 det ( x exp
~]02R~
F
[~~o~r i. 'x" , x'., T )  i 2 D u ( T ) ]
,
(3.2.24)
where Hamilton's principal function is given by R o~r , ' x " x, ' T;'  )
raw [ (x'2+ x,,2) c ~ 2sinwT
]
(3.2.25)
3.3 T h e Radial H a r m o n i c Oscillator
Let us now discuss the most important application of the radial path integral (2.7.15), namely the radial harmonic oscillator with V(r) = 89 2. The original calculation is due to Peak and Inomata [771], and has been investigated by Duru [274], Goovaerts [402], and Inomata [518]. We present the more general case with timedependent frequency w(t) following Goovaerts [402]. We have to study the kernel r(t")=~"
K,(r",r';t",t') =
f
Dr(t)pt+p.~[r2]exp
r(t')=r'
= lim(
m
drnpl+ D;2 [r 2]
• exp [h 2 ~ = ~
lim N+co
e i~(r'~+r''~)/2
!~__ fl" h 2 Jr, (§
_ ~2(t)r2)d t
]
100
Basic Path Integrals
x 1"I L j=l
rjdrjexp [i(fl, r~
+/7,r~ +...+ flN_lr~/_l)]
x [el+_~ ( i OtrOrl) X . . . x ll+ D,a ( i o~rN_ l rN) ]
(3.3.1)
(~ = m/eh, flj = a(1  e2wy/2)). By means of the integral formula [413, p. 7181 L ~176 xe  ' ~ J.(~x)Jv(nz)dx = ~1 e(~=+'a)/4' 1~ (cJ~yfl)
(3.3.2)
and its analytic continuation to pure imaginary argument [771], the convoluted integrations are performed through a recursion and it is found that the kernel IQ can be cast into the form
K,(r",~';t",t') = ~vSW
m ibm(T)
rim f~(T) ,,
x exp [~" t ~   ~ r
, I t ) ,,=~1 )j/i+~g~(\ ~~.,.,, ]
(3.3.3)
+~~
The quantities rl(T) and ((T), respectively, are determined by the following differential equations with boundary conditions /j+w2(t)rl = 0 ,
q(t') = 0 ,
i/(t') = 1 ,
(3.3.4)
~'+w2(t)~ = 0 ,
{(t') = 1 ,
~(t') = 0 .
(3.3.5)
In particular for w(t) = w (time independent) we obtain 1
o(t) =  sinw(t  t') ,
r
= cosw(t  t') ,
~(t) = cos~(t  t') .
~0
This yields the radial path integral solution for the radial harmonic oscillator with timeindependent frequency
Kl(r", r' ; T) = ~
x oxp
mw
[
i h sin wT ~rR,o1,tr//
9
(ihsin
,
. ,:,.3.~
The next step is to calculate with the help of (3.3.6) the energylevels and wave functions. For this purpose we use the HilleHardy formula [303, Vol. II, p. 189] (Itl < 1) t~12
1
,l+t]
1t 
~
tnn!e89
n=0 F ( n + a + 1)(xY)~'/=L(a)(x)L(~a)(Y)
(3.3.7)
3.3 The Radial Harmonic Oscillator
101
where L (~) are Laguerre polynomials. With the substitution t = e 2iwT, x = mwr'2/h and y = mwr"2/h in (3.3.6) we get finally: oo
K,(r",r';T) = Z eiTEN/h R~(r')R~(r")
(3.3.8)
N=0
X exp (~r) rn~
2
/.(/+~) \ n r 2 ]
 ~.:j.
.
(3.3.10)
The path integral for the harmonic oscillator suggests a generalization in the index I. For this purpose we consider equation (2.7.15) for D = 3 with the functional weight/~t+ 89[r2]
,(t")=r
/
ri
t"
]
..(,,.,+,,..,ex.
9
r(t')=r'
(3.3.11)
The functional weight corresponds to the centrifugal potential ~ in the Schrhdinger equation. Replacing l + 1 by A corresponds to the "centrifugal potential" Vx(r) := h 2 (A2  88 2 in the Schrhdinger equation where A can also be complex, provided ~A > 88 This situation corresponds to the classical radial Lagrangian s = ~§  h 2 ( A 2  88 2  V(r). With Y(r) = ~ 2 r ~ this leads to the path integral expression r
 h2 2mr 2
dt
Q
(3.3.12)
r(t')mr'
Let us remark that in the literature one often uses the asymptotic form of the modified Bessel function, e.g. [771]
I~(z) ~_ (2~z)89188
(]z] >> 1,
~(z) > 0) ,
(3.3.13)
and the functional weight becomes (ignoring the condition ~(z) > 0): p~[r=] (,~0) i N ~ exp {  h ~ e h2(A~88
.
(3.3.14)
According to Fischer, Leschke and Mfiller [343] the functional weight approach according to (3.3.11) defines expressions such as (3.3.12) in a unique
102
Basic Path Integrals
way such that we can use the asymptotic form of the functional weight in the path integral still and thus have the following equivalence in terms of the functional weight formulation (for more details we refer to the literature, e.g. [343,464, 528, 771,865])
r(t")=r" r(t')=r' __
/
~r(Q#x[r2]
(§ _~o2r2)at
r(t')=r'
 ihsin~T
exp t~~r
tr"2) cot~T
]
I~ \i~~sin~T] .
(3.3.15)
This path integral identity is the second of our basic path integrals, the
Besselian path integral. 3.4 Path Integration Over Group Manifolds Before we discuss in detail the path integration on specific group manifolds, let us start with some general remarks. We are interested in a general formalism which tells us how to perform path integration on a group manifold. However, we must first ask, what are the relevant objects in the group manifold we want to concentrate on? Is it possible to perform the group path integration in such a way that we can calculate the kernel explicitly? Is it more appropriate to consider an expansion into the characters of the group, respectively into matrix elements in the coordinate space representation? As we will see, it depends on the group which of the alternatives can be applied. Actually, most authors have concentrated on a character expansion of the path integral, see e.g. [104, 262, 263, 679,726, 776]. However, this is not the only possibility. It has the advantage to be coordinate independent, though. For several applications, an expansion into the matrix elements in the coordinate space representation is nevertheless desirable. Path integral identities can be extracted from such spectral expansions. In some cases it is also possible to state the kernel explicitly, and in other cases the Green function can be explicitly stated. In the following we present some general arguments that tell us which calculational tools can be used to consider and perform path integration on a group manifold. In our presentation we follow the lines of reasoning of BShm and Junker [104] (see also [106,523,551]). First we will give an introduction to the general theory, followed by the path integration over SU(2) and SU(1, 1). These two group manifolds and the path integral identities derived from them will be of great importance in the applications, where we have to deal with P5schlTeller potential and modified PSschlTeller potential problems, respectively.
3.4 Path Integration Over Group Manifolds
103
3.4.1 G e n e r a l F o r m a l i s m . In order to set up our notation we start with a Ddimensional flat space with an indefinite metric according to
(gab) = diag(~l,...,+l,
cl,...,1)
p times
.
(3.4.1)
q times
Consequently, we write a classical Lagrangian on the lattice for such an indefinite metric as follows =
"'
)
(3.4.2) In order to take into account the indefinite metric in this pseudoEuclidean space Ev, q endowed with the metric (3.4.1), we must change the measure in the path integral according to K(x",x';T)
x exp
=
lim /
m
~NP/2(irn)gq/2N~/~
N>~ t,27TTg~h)
_l h j1 ~m =
~
=
(Azk)2  E (Azk)2 / k=p+l /
,+,
dxt
~'V(Xj)
(3.4.3)
To regularize such a path integral, the various integrals must be treated separately [104]: in the integrations over the variables with positive signature (compact variables) in the metric, a small positive imaginary part has to be added to the mass (m + m + i rl, r/ > 0), whereas in the' variables with negative signature (noncompact variables) a small negative imaginary part (7/+ m  iT#, r / > 0) must be added. We consider the scalar product (x,x) = ( z l ) 2 + . . . + ( z P )
2  (zP+l) 2  . . . 
(xP+q) 2 ,
(3.4.4)
and introduce the sets
T+I = {x [(x, x) > 0} T_~ = {x I(x, x) < 0}
(timelike)
t
(spacelike) .
J"
(3.4.5)
We want to achieve in the following an expansion of exp[z(ej_l, ej)] in terms of the matrix elements of the corresponding group representations (in particular for rotation groups). Let G be such a group and g 6 G. Let s be a linear vector space of functions f, actually s = s Then f(x) E/Z2 +_~f(gx) E/:2 ,
x E I R p+q
for any g 6 G. A regular representation of G is given by
D(g)f(x) = f ( g  l x )
.
(3.4.6)
104
Basic Path Integrals
Such a representation D is decomposable into a discrete sum of unitary irreducible representations D t in the Hilbert space s Furthermore, let H be a subgroup of G which leaves the nonzero vector a E L:2 invariant, i.e.,
D r(h)a = a,
hEH C G .
(3.4.7)
Let (G, H) be a Gelfand pair, then G / H is a symmetric space and
s
= ( ~ Dt, lEA
(3.4.8)
where l E A denotes the class one representations. With each vector f E/~2 we may associate a scalar function
f (g) = (D l(g)f, a) .
(3.4.9)
The f (g) are called spherical functions of the representation D l (g). Choosing a basis {bi} in s so that b0 = a, the matrix elements of Dr(g) are given by
D~n (g) = (D r (g)bm, b,~) .
(3.4.10)
In particular the Dl0,~ are called associated spherical functions and the Dr00 the zonal spherical functions [911]. We have the identities (h E H)
D~o(gh) = D~o(g ) ,
Dtoo(h'gh) = Dloo(g) .
(3.4.11)
An important property of the spherical functions is that they are eigenfunctions of the corresponding LaplaceBeltrami operator ALB on the homogeneous space M = G / H . Of course, the Hilbert space s is spanned by a complete set of associated spherical functions Dt0m. We introduce harmonic analysis, i.e., a generalized Fourier transformation on locally compact groups: let f E s and D,~ l n as defined in (3.4.10). The Fourier transform ]~n of f(g) is defined as f(g)=
^L dEtdtEf~,D , n , ~1 ( g ) ,
(3.4.I2)
rrl,rl
i~,~ = fG f(g)D~* (gZ) dg ,
(3.4.13)
and dg is the invariant group (Haar) measure, f dEt stands for a LebesgueStieltjes integral to include discrete (fdEz .+ ~ t ) as well as continuous representations, f dEt is to be taken over the complete set {l} of class one representations, dl denotes (in the compact case) the dimension of the representation and we take
dt fG Dmn(g)Dm,n,(g)dg = J(l,t')g,~,m,J,~,,~, 1 l' *
(3.4.14)
3.4 Path Integration Over Group Manifolds
105
as a definition for d~. 5(1,1') can denote a Kronecker delta, respectively, a (ffunction, depending on whether the quantity l is a discrete or continuous parameter. We have furthermore the group composition law l*
l
(3.4.15)
Dlmn(galgb) = E Dkn(ga)Dkm(gb) "
k
We must analyze under which conditions (ej_ 1, ej) can be expressed in terms of the group elements gj1 and gj in order to apply the expansion (3.4.15). Let G be a transformation group of Ha, i.e., e:ga,
e, a E Ha
(3.4.16)
where g E G denotes a (D x D) matrix representation of G (D = p + q), and e, a E Ta with the same ot = 41. The unit sphere Ha is covered by all possible rotations. Possible choices for G are groups which contain SO(p, q), respectively SU(u, v). For example, the fourdimensional sphere S (3) is isomorphic to the group manifold SU(2) and instead of SO(4) we may choose SU(2) as a transformation group of S (3) = SO(4)/SO(3), i.e., a rotation group. Generally, we are most interested in a path integral representation in generalized spherical coordinates in some geometry in the sets Ta E Ev, q, i.e., x = rH
re(O 1, . . . , 0 p+q1 ) ,
v:
1,...,p+q
,
(3.4.17)
where the e are unit vectors in Ta. The e span the unit sphere Ha: Ha = {e[(e, e) : a}, a : + 1 ,  1 . We express the Lagrangian (3.4.2) restricted to Ta in terms of the spherical coordinates (3.4.17) yielding (x E lKv+q)
=
m 2 "4~2 {(Arj) '~ 2rj_lr j [1 q= ( e j  l , e j ) ] }  V(xj)  AV(xj)
m 2 : t~e2 [rj_ 1 + r~ ={=2rjlrj(ej1, ej)]  V(xj)  AV(xj) .
(3.4.18)
Therefore we obtain the path integral representation
iqN/2N_lro~ N~oo \21rich] "
hj=l
(
m
2
~
~+qldrj m
/ dH~
106
Basic Path Integrals
In (3.4.18) and (3.4.19) AV denotes a quantum potential constructed in the usual way in terms of the metric, but with the modification that the kinetic terms of the Lagrangian are rewritten in terms of (ej1, ej) which requires a carefully calculated Taylor expansion (compare the discussion of the sphere and the hyperboloid). Although a path integral representation in spherical coordinates may be in most cases sufficient and convenient, it is not the only possibility. In particular in the case of homogeneous spaces, the corresponding path integral allows as many spectral expansions in coordinate space representations as there are separable coordinate systems in this space. The path integral representations in spherical coordinates have allowed the derivation of several of the basic path integrals, among them the path integral of the radial harmonic oscillator, the PSschlTeller and the modified PSschlTeller potential (see below in the discussion of the path integration on the SU (2) and SU (1, 1) group manifolds). In Sect. 3.4.2 we discuss interbases expansions which allow us to switch in the path integral from one coordinate space representation to another. These new coordinate space representations of a path integral in a homogeneous space can then give rise to new and more complicated path integral identities. The method has been extensively worked out in [447], and several expansions are listed in the table of path integrals. In the following we concentrate on two principal possibilities of harmonic analysis of a path integral on a group space. They are i) Ha is isomorphic to the group manifold G: Ha ~ G. H~ ~_ G is quite a strong constraint, so it is not surprising that there is only a very limited number of groups which satisfy it. The harmonic analysis in these cases is performed by the characters of the group Xl (g). In the following table the appropriate requirements are listed, i.e., dim Ha = dim G. Table3.1. Dimensions of homogeneous spaces G
dim G
SO(p,q)
(p+q)(p+q1)/2
su(u, v)
(u + v) ~
dim H,,
dim G = dim Ha
p+q1
p+q=2
2(u+v) 1
u+v=2
We see that only the four cases ]SO(2),
SO(1,1),
SU(2),
SU(1,1)]
remain. For the oneparameter groups SO(2) and SO(l, 1) the irreducible representations are onedimensional and in fact the general Fourier transformation (3.4.13) is reduced to the usual Fourier and Laplace transformation, respectively. Therefore we are left as the only nontrivial examples with SU(2) and SU(1, 1), which we discuss in two subsequent sections.
3.4 Path Integration Over Group Manifolds
107
ii) Ha is given by a group quotient: Ha = G/H. This case describes motion on quotient space group manifolds, i.e., on a homogeneous space. Examples are the motion in (pseudo) Euclidean spaces, and on spheres and hyperboloids. However, the method is more general and it is not restricted to these two cases. The harmonic analysis in this case is performed by means of the zonal spherical harmonics ntoo(g) [911]. 3.4.2 I n t e r b a s e s E x p a n s i o n s . Other important tools in group path integrations can be derived by interbases expansions, i.e., for problems which are separable in more than one coordinate system. In the case of potential problems, these potentials are called superintegrable. This property is very closely related to the fact that such problems have several integrals of motion, and the underlying dynamical symmetry group allows the representation of the problem in several coordinate space representations of the group. Superintegrable systems can be found in Euclidean space [305, 458, 668], in spaces of constant curvature [457462], and in the theory of monopoles [461]. The basic formula is quite simple, being ik) = / dEt Cl,kll) ,
(3.4.20)
where Ik) stands for the momentum space representation of the eigenfunctions with quantum numbers k, and f dE1 is the expansion with respect to 1 with coefficients Cl,k which can be discrete, continuous or both. The difficulty is, provided that one has two momentum space representations in the quantum numbers k and 1, respectively, to find the expansion coefficients. The expansions which involve Cartesian coordinates and spherical coordinates are well known. For the problem of the free quantum motion in Euclidean space, this means that exponentials representing plane waves are expanded in terms of Bessel functions and circular (polar) waves (a discrete interbases expansion). This expansion actually is the very starting point in the formulation of radial path integrals, and the path integral evaluation of the radial harmonic oscillator, see Sect. 2.7. Because the expansion coefficients are unitary, i.e. (Cl,k) 1 = C~,k, we can insert them twice in the spectral expansion of the Feynman kernel and obtain the identity / d E k ~k(x")kV;(x') e iEkT/n = / dElr
iE'T/r' , (3.4.21)
where we have chosen the coordinate space representation of the wave functions with coordinates x and quantum numbers k, respectively coordinates u and quantum numbers 1, representing two equivalent solutions of the same problem. This general method of changing a coordinate basis in quantum mechanics can now be used in the path integral. We assume that we can expand the shorttime kernel, respectively the exponential e ~x~~Xj in terms of matrix
108
Basic Path Integrals
elements of a group in the desired coordinate space representation. We can then change the coordinate basis by means of (3.4.21). Due to the unitarity of the expansion coefficients Cl,k the shorttime kernel is expanded in the new coordinate basis, and the orthonormality of the basis allows us to perform explicitly the path integral, exactly in the same way as in the original coordinate basis. The use of interbases expansions for the calculation of path integrals has been indispensable in almost all cases of the nontrivial basic path integrals. At the basis for the evaluation of the radial path integral are the interbases expansions (2.7.12), which expand a plane wave into Bessel functions. Interbases expansions must also be used to evaluate path integrals in other coordinate systems, like elliptical and spheroidal coordinate systems, see e.g. [447] for many examples. 3.4.3 P a t h I n t e g r a l s o n H o m o g e n e o u s Spaces. Let us consider a path integral representation on a homogeneous space ]M [519, 523, 911]. Let a group G be a transformation on this space IM. If G acts transitively on IM, then IM is a homogeneous space with respect to G. In order to convert a path integral on a homogeneous space into one on the corresponding group manifold, we require that the shorttime kernel has some invariance properties: We assume that the shorttime kernel is symmetric under the interchange of the two end points q' and q" (q', q" E IM, the space lM is allowed to have an indefinite metric), and is invariant under the action of g E G, i.e.,
K(qj, qj1; e) = K ( q j _ l , qj; e) = g ( g qj,gqj1; e)
(3.4.22)
for all g E G. For an arbitrary fixed point q~ we introduce the function
k(g; e) = K(g qa, qa; e) .
(3.4.23)
K(qj, qjz; e) = k(glg'; e) = k(g'lg; e)
(3.4.24)
We then derive
with qj = gqa, qj1 = g' qa. Consequently, we can define the Feynman path integral as the limit of a multiconvolution (denoted by *) N
K(q",q';T) = lim ][ *k(gT~lgj) , N ~ c o , L . L
j=l
(3.4.25)
J
where qj = gj qa, j = 0,..., N. Thus, the path integral in a homogeneous space is reduced to a convolution on a group manifold G. Due to the property of a homogeneous space H~  G/H, we know that the harmonic analysis can be performed by the zonal spherical harmonics, i.e., we can expand by using the subgroup composition g = ab with a E A, b E B such that G = A  B according to ()~t := ]too)
3.4 Path Integration Over Group Manifolds
109
k(ab; e) = / dEl d, Xt(b; e)Dloo(a) .
(3.4.28)
The expansion coefficients in turn are given by
At(b; e) = JfA da k(ab; e)Dtoo(a 1) .
(3.4.27)
This expansion allows us to perform the convolution on the group manifold G yielding g(q",q';T) =/dEj J
dtgt(b'lb";T)Dtoo(a'la" ) ,
(3.4.28)
where N
Kt(b'lb";T) = N~ lim H A'(bf ~lbj;e) " oo
(3.4.29)
j=l
A special case can be stated if the subgroup B consists of the unit element e alone. Then Al (e, e) = Xl (e), and we can explicitly state the corresponding limit yielding the expansion K(q,,q~;T)
= _/dEldtes
I ,1 a ,~) Doo(a
/ dEt Z eiE'T/h(q" Ilrn)(lmlq') ,
(3.4.30) (3.4.31)
171
(q
Ilm) = V/~t D2o(a) , El = i hit (0) .
(3.4.32) (3.4.33)
Let us emphasize again that the index l can be discrete as well as continuous. If the homogeneous space is compact, l is a discrete quantum number. Examples are the path integral representation on the sphere or path integral representations in compact subspaces of noncompact spaces. In noncompact spaces such as the hyperboloid A o1 and (pseudo) Euclidean space, l is a continuous parameter.
3.4.4 The SU(2) Path Integral.
3.4.4.1 Cylindrical Coordinates. The group manifold SU(2) is of particular interest because it serves as a model for spin. Considering a quantum mechanical spherical top one can distinguish an external motion and an internal motion. Its path integral can be separated in terms of these two independent motions and it is not required that an interaction has to be turned on. The internal motion is now interpreted as "spin" taking on integer as well as half integer values. The first discussion of motion on a group manifold was due to Schulman [826] in his discussion of the spherical top in terms of Euler angles.
110
Basic Path Integrals
Schulman made an analysis comparing different approaches, i.e., as seen as motion on a curved manifold and by exact summation of the classical paths using the semiclassical approximation. In this particular case, the dimension of the group manifold is isomorphic to the covering unit sphere 7{a, e.g., we have SU(2) ~_ S (3) . Schulman was interested in constructing a path integral for spin, and in fact in the matrix elements of the (2J + 1)dimensional unitary irreducible representation of the Hilbert space of SU(2) the "angular" momentum can take on integer as well as halfinteger values. The eigenvalues for the corresponding Casimir operator take on the values 1 , 1,3 ~, ...). However, since SU(2) "" S (3), we can look Lj = J(J + I) (J = 0, ~ at this group manifold from the point of view of motion on the sphere S (3) . On the S D 1_sphere in turn the eigenvalues of the LaplaceBeltrami operator are given by Lt  l(l+D2), thus we have for the S(3)sphere L~ = / ( / + 2 ) = ~1L J, where we identified I = 2J and the halfinteger representations are "hidden". In order to derive the correct Feynman kernel for the SU(2) path integral we start by considering a specific coordinate representation. This enables us to determine the correct quantum potential to be included in the effective Lagrangian. Rewriting the coordinate representations in terms of Tr(galgb) gives us finally the appropriate form to deal with the various expansions. Motion on SU(2) means that the quantum mechanical motion is subject to the constraint x~ + x~ + x32 + x42 = 1 , (3.4.34) i.e., it takes place on the S(3)sphere. A suitable coordinate representation, called cylindrical, has the form (0 _< ~ < 2r, 0 < 0 < rr, 0 _< r < 47r) o
xz=sin~sin o.
2
,
x2=sin~cos
~+r
X3 ~ COS ~ sin
o
2
'
2
(3.4.35)
~+r
x4 ~ cos ~ cos
2
The classical Lagrangian reads as m . :1 + ~ + ~2 + x4~) __ /:(x, x)  ~(x
2
+
+ 2~br c o s tg)
(!o
(3.4.36)
We read off the metric tensor and its inverse, respectively
1( o
(gab) ~
0 0
1 CO 0 cos t9
,
(gab)=4
sin 20 cos 0 sin 2 0
0 COS
sin 2 ,9 1
sins 0(3.4.37)
and g = det(gab) = (sin 0/8) 2. We have F~ = Fr = 0 , Fd  cot ~, and for the momentum operators we obtain Pv 
i (9~ '
PC = ]~~ ,
Pd = ~ ~
+ 
(3.4.38)
3.4 Path Integration Over Group Manifolds
111
The quantum potential AV is found to be ZWwoy~(~) = ZWpF(*3)  ZW(*3) = ~~ 1 + ~
.
(3.4.39)
Thus we have the necessary ingredients to write down the path integral for quantum motion on SU(2) in terms of Euler angles o(t")=,~"
f
O(t~)=O '
x exp { i'f t : "h [ 8 (*32+ tb2 + ,2 + 2~br cos ,3) + ~mh2( 1 + ~ 1
) ] dt }
o
(3.4.40)
On the other hand we know that the SU(2) path integration can be performed by using the character expansion in the harmonic analysis. We thus obtain for the SU(2) path integral in the character expansion Xt = C~j(cos ~) together with the corresponding spectral expansion KSV(~)(0", r = ~
~", .3', ~', r T)
2J+1 1 (_~) 2r 2 C2j cos
exp
[
2i
]
mtiTJ(J+ 1)
(3.4.41)
J=0,89
%.nntv ,~o ,.,. ,_mn,v ,~o',r
=
2J(2J+2)
J,"`,n
(3.4.42)
where *3,
*3,,
cos yI2 __sin ~ sin T cos
~,, _ ~, _ (r
_
r
2
v9' cos y~" cos ~o"  r + 2(r  r + cos ~
'
(3.4.43)
and with the wave functions a
~
1
eirn~o+inr
J
(cos*3)
=2" /2J+11'(grm+1)r(g+m+1) V ~
eim~o+in r
~ ( ( g  n 7 1 ) F ( J + n'+ 1~
~ ~ n("`n,"`Tn) • ( 1  cos *3)~~"(l + cos tD ~ t)_ m
t
(cos*3) , (3.4.44)
112
Basic P a t h Integrals
where the P(na'b)(x) are Jacobi polynomials [413], and the energy spectrum is h2
E j  ~ m 2 J ( 2 J + 2) . The e(a,b)and the DJn are related by (s = J [ m + re'l, Wigner polynomials) [908]
=
(3.4.45)
1(# + v),/J = I m 
m'h
I s!(s+#+v)!
7 .(7 ,,u~ff';~]!(1  z)#'/2(1 + z)~'12P("'~')(z)
=
9
(3.4.46) Note that 2 J E IN and thus the energy spectrum is indistinguishable from the motion on the covering S(3)sphere, as it should be. Note that due to the fact that the group manifold SU(2) is isomorphic to the S(a)sphere, there are six coordinate space representations of the SU(2) path integral, i.e., as many as there are separable coordinate systems for the LaplaceBeltrami operator on the sphere S (3). Therefore the character Xl may be expanded into six different coordinate space representations of the wave functions. The switching between the various spectral expansions is performed by the relevant interbases expansions. The combinations of corresponding path integral representations and spectral expansions yield path integral identities. In particular, the spherical coordinate system on the S (3)sphere (the threedimensional rotator) yields a path integral identity for the 1/sin20 potential, a special case of the path integral identity of the PSschlTeller potential.
3.4.4.2 The PSschlTeller Path Integral. Let us consider the (PSschlTeller) potential
+
cos
.
'
0
0)
118
Basic Path Integrals r(t")=r"
h ~ ~ 1 7 6 eiET/h
/ Dr(t) r(t')mr'
m V(ml  Lv)l'(Lv + ml + 1) li2 F(ml + m2 + 1 ) F ( m l  m2 + 1) x (cosh r)("~m~) ( tanh r>) m,+,~+ {
(3.4.69) 3.4.5.3 Horocyclic Coordinates and the Inverted Liouville Problem. Let us turn to the horocyclic system. It parametrizes the coordinates on the hyperboloid (3.4.52) as follows (v = (v0, vl) 9 IR(I'U, u 9 IR) xl = 89 = +e=(1  v2)] , xu=v0e u ,
x3 = Yl eu , x 4 = ~1 [  e   u + e * ' ( l + v 2 ) ] ,
(3.4.70)
with the domain of the coordinates as indicated above. As known from the theory of harmonic analysis on this manifold one has a discrete and a continuous series in the variable u with wave functions corresponding to (k = (k0, kl) 9 IR(1'1), k 2 = k02  kl2 > 0 is taken in the physical domain, Ji k (z) = [Ji k (z) + J_ i k (z)], a is the parameter of the nonunique selfadjoint extension) [559, 696, 893] Discrete series (n E IN, 0 < a < 2):
~'fk~
e i kovoi 
klvl
271"
X/2(2n+~) J2"+~(lkle")
'
(3.4.71) Continuous series (I k l > 0):
eik~176 ~Pko,hl,h (v0, Vl, u) 
2~r
i
k 2 sinh irk
&h(Ikle")
9
(3.4.72)
These wave functions form the matrix element expansion of the Titchmarch transformation. According to the general theory of the path integration on group manifolds we have to calculate the quantity Since this expression is actually independent of the representation one chooses we can take the
3.4 Path Integration Over Group Manifolds
119
result of BShm and Junker [104] and we have in the limit e + 0 the result (] k ]2 = k02 _ k~ > 0, i.e., k in the physical domain) Discrete series (n E gq): h"sUOJ) = exp
4n 2  1 (3.4.73)
Continuous series (p > 0): u sU(1,1) ~* 89

exp
(
ihT'2 )  ~m (k I 1)
(3.4.74)
Putting everything together we arrive at the following path integral representation
=(t")==" v(t")=v" [ i t" / "Du(t)e2u / ~Dv(t)exP[hftl (2(h2e2~'~2)2h~2m)dt] ~(t')=~,'
e~'+~'' [ x
v(t')=v' e i k .(v"  v') dk (~,~) 47r2 2(2n + c~)J2,+~(I k le
)J2n+a(IkleU')e ih(4n21)T/:tm
I. nEIN
kdk
~k(Ikhe2 sinh irk '
+
)Jik(Ik[e ~' )e ih(k%l)T/~'~l (3.4.75)
This result enables us to derive the path integral identity for the inverted Liouville problem. Separating off the (v0, Vl)path integrations we obtain y(t")=y"
y(t')=y'
=
2(2. +
e 'ihT"
/m
II
yl
)
nfi~l
+
fo ~ ) si~rk kdk Jik ~ (to e y , )Ji~k(tcey,,)e_ir, k2T/2m
(3.4.76)
3.4.6 E x p a n s i o n Formulae. In this section we give a short list of expansion formulm which emerge from a particular coordinate representation of the Euclidean group, say in IR 2 and IR3, and in pseudoEuclidean spaces or hyperboloids, respectively. The expansion of plane waves in spherical waves (2.7.12) has already been exploited in Sect. 2.7.
3.4.6.1 Elliptic and Spheroidal Coordinates in Flat Space. As has been shown in [447] expansion theorems for elliptic and spheroidal coordinates
Basic Path Integrals
120
can be used to derive explicit path integral representations in these coordinates; the expansion formulae are in fact interbases expansions. Let us start with the twodimensional case. We consider for an arbitrary a the periodic and nonperiodic Mathieu functions men, Me (1), and the corresponding even and odd Mathieu functions can, sen, Mc (1) , Ms (1). We have the relations Cen(Z; h') = men(z; h')/vf2, MO)(z) = McO)(z; h) (n = O, 1,...) and sen(z; h') = i. men(z; h')/v/2, M(')~(z) = (1) n Ms(1)(z; h) (n = 1, 2 , . . . , h = kd/2, k is the wave number, d is the distance between the foci of the ellipse). The orthonormality relations are given by [692] 1
ce.(O) cer(~)d0 = :
sen(0) set(0) d~
= 
me~(0) me~(O) d0 = ~nt ,
2~"
1
[" een (0) set (o) dO = 0
(3.4.77)
7r
.
(3.4.78)
d'~ocod # f;dv(sinh'/.z 4sin'u)Mn(l)(/.z; h)M~ 1)*(jU; h')men(t); h')me~(t);h") = 2[(fnt6(k  k') .
(3.4.79)
We have the following expansion of plane waves into elliptical waves [692, p. 185] oo
exp [ik(x cos a + ysin a)] = 2 Z
in cen(a; hS)MO)(/~; h)cen(v; h 2)
n:0 OO
+ 2 ~ i n sen(a; hS)M(J,)(~u; h)sen(u; h') . .=: (3.4.80/ In the limit d + 0 the functions M (1) yield JnBessel functions, and men exponentials. In three dimensions one has for prolatespheroidal coordinates the expansion [692, p. 315] in spheroidal wave functions S~ (:)(z; kd), ps~ (z; kSd ') exp [i kd(sinh/~ sin v sin 0 cos ~o+ eosh ~ucos v cos 0)] co
=~
l
Z(2/+l)
i '+Sn
1=0 n=I
x S;' (1)(cosh/~; kd) ps? (cos v; k'd 2) ps7 n (cos O; k2d 2) e i nv (3.4.81) The orthonormality relations are given by
3.4 Path Integration Over Group Manifolds d3
f0 /0 dp
dv(sinh 2 ~u + sin S u)
f0
121
d~o e i so(nn')
x S t (1)(cosh #; kd)S~n,' (1)(cosh #; k'd) ps• (cos v; k2d 2) ps vn t (cos v; k'2d 2) 2 (l + ")! = 21 +~(_S~).t
k
(3.4.82)
 k')
f0r sin t9 d~ f02r d~o psln(cos~;k2d2) PSvn'*(cosd;k2dg)eiso(nn') 47r (I  n)!~nr,,~u, . : 2t +i
(3.4.83)
In the limit d + 0 the functions S/'(1) yield Jl+l/2Bessel functions and the ps~n Legendre polynomials, respectively. The case of oblate spheroidal coordinates is similar and follows from analytic continuation, cf. [447].
3.~.6.2 Spherical Coordinates in PseudoEuclidean Space. The expansion formula for (pseudo)spherical coordinates on the hyperboloid has been important in the evaluation of the path integral on the hyperboloid, cf. [104,466]. Actually, it describes the interbases expansion of plane waves in a space with a Minkowski spacetime metric into spherical waves. It has for some a the form [413, p. 804] e~ cosh ot =
(Z sinh a) '3
F(ik)T
79i~89(c~ (3.4.84)
which for D = 2 takes on the form eZ cosha _ l~r/1~,dk eika Kik(z) 9
(3.4.85)
K~,(z) is the MacDonald Bessel function, and P~(z) is the associated Legendre function. This formula is valid in D dimensions. For more details, cf. [447]. 3.4.6.3 Elliptic and Spheroidal Coordinates in PseudoEuclidean Space. Similarly to Euclidean space, expansion formul0e in elliptic and spheroidal coordinates can be found which have the correct limit for d 4 0. Me,(z; d 2) or e "" and M(3)(z/d; d) oc H(1)(z) (d 4 0), and they yield the wave functions of the polar system. They obey the orthonormality relations (h = kd/2) 1
f drMeik(r;h 2) Me~k,(r; h ~) = J(k' 
k)
(3.4.86)
P eirk 2 [ /'~176 d L da./, db (sinh2 a  s i n h 2 b ) x Mei k(b; ~ ) M e ; k , ( b ;
= ~ ( k '  k)~(p'p) .
e_~_~4J'"ikA/r(3),~, r ~)Mi(~r !* (a; P~) (3.4.87)
122
Basic Path Integrals
Therefore we have the expansions for plane waves in the coordinates (v0, Vl) with metric ( + ,  ) exp [ip(v0 cosh r  vl sinh r)] 1 J dk e  ' k / 2 Ueik(b; ~4e~) Mei ~(r; e2_~ a/r(3)(a; v_4 4 ]"~ik 2)
~" 2
"
(3.4.88)
In three dimensions matters are a little more complicated. We must extend the expansion of the twodimensional case to three dimensions with the restriction that for d + 0 we get back the spherical system. The proper spheroidal functions consequently are [692] Psi(z; 72) and Sgik1/2((3) 'Z'/%" ")'). S ~ (3) They have the asymptotic behaviour P s i ( z ;'y2) o( P~ (z) and i k1/2('z"/'Y; 7)'
o~ ~ H~lk)(z) ('7 =pd, d ~ 0), giving the spherical wave functions. From these considerations we derive by using the theory of spheroidal functions [692] the following interbases expansion [447] exp [i pd(cosh ~ cosh ~/cosh a  sinh ~ sinh 71sinh a cos ~)]
1
:
dkksinh~rklF( 89
7r3/2
+n)12e'~k/2e inv
• PsV". 1 (cosh,; p2d ) Ps72 1/2(cosh.; p2d2)S ' / (cosh
pd) (3.4.89)
The spheroidal wave functions obey the orthonormality relations de
sinhada
sik_l/2(cosh a) Ps~s
a) e i~(nn')
2r2 ksinhTrk[F(89 + i k + n)l~5,,,~i(k  k') ,

d3
d~
(3.4.90)
dr] (sinh 2 ~  sinh 2 y) sinh ~ sinh rI
d~ e i ~(n#) dO
• Ps~k~ 1 ~(cosh 7};p2d2) Ps~k~ 1/2(cosh 7};p'2d2) , . . ~ (3) ...... ,~' ( 3 ) . , . • ~i k 1/~1,c~ r Pa)~i k,1/2t c~ 5; p'd) 71"
 ksiZ~r ~' 
1
 2 27i"
~e
r~k
5,~,vS(kk'),i(pp')
.
The case of oblate spheroidal coordinates is, of course, similar.
(3.4.91)
4 Perturbation
Theory
In this chapter we study perturbation theory within the path integrals. We summarize four approaches for the evaluation of perturbation expansions which are: i) Path integration of a perturbation expansion after the method of Devreese and Goovaerts. In this method one expands the potential term in the action in a perturbation expansion, performs a Laplace transformation of the potential (if possible), and reduces therefore the original path integral problem to that of a specific quadratic Lagrangian. The latter path integral problem can be solved exactly and one is left with a sum in powers of the coupling constant for the (energy dependent) Green function of the original potential problem. ii) Timeordered perturbation expansion of the path integral and boundary problems in path integrals. This expansion method is the original perturbative approach: if it is not possible to obtain for a quantum mechanical problem an analytical expression of the kernel, respectively its Green function, one expands the Feynman path integral about a known solution in powers of the coupling. This requires that the coupling constants are "small" in order that the perturbation expansion makes sense. Usually, it is not possible to perform all the integrations involved except for specific problems. The most important among these specific problems is point interactions which are treated in some detail. They belong to the class of exactly solvable quantum mechanical problems, and lead also to the incorporation of explicit boundary conditions in the path integral. iii) Effective potentials in partition functions. In many applications in statistical physics approximate solutions of the generating functional are sufficient for a numerical investigation. Instead of exactly solving a path integral problem, one is interested in a "good and fast" algorithm. The theory of the effective potential in partition functions provides such an algorithm. It exploits the solution of the harmonic oscillator in order to derive a prescription for the numerical determination of an effective potential. iv) The semiclassical expansion about the harmonic expansion. The semiclassical expansion makes full use of the information contained in a general quadratic Lagrangian, respectively Hamiltonian. It allows one to expand an arbitrary dynamical quantum mechanical problem about its semiclassical solution in powers of h up to any order. The expansion in powers
124
Perturbation Theory of h is based on moment formulae which are the quantum mechanical analogue of Wick's theorem.
4.1 P a t h Integration and Perturbation Theory In this section we present the perturbative method for path integrals for quite arbitrary potentials. The path integral is rewritten in terms of a specific quadratic Lagrangian for which the path integral can be solved exactly. Finally, one is left with a perturbation expansion in powers of the coupling constant of the potential. The method was originally developed by Feynman [340] and Devreese et al. [405, 408, 409] with the explicit treatment of the Coulomb potential. Further examples are the usual harmonic oscillator treated by Grosjean and Goovaerts [472] (slightly modified) and the Jfunction potential by Goovaerts, Babceno and Devreese [404]. The drawback of the method is that it is quite involved for all the standard examples such as the harmonic oscillator, the Coulomb potential, or the 5function. The case of the harmonic oscillator leads to a very cumbersome perturbation expansion, and in the case of the Coulomb potential, only the wave functions of a definite parity could be extracted. Also, the principal problem that the kernel of the Coulomb potential is not known in closed form cannot be resolved. However, this method showed for the first time that it was possible to extract all the relevant information of the discrete and continuous spectrum of the Coulomb potential by a genuine path integral approach. 1 The first object which was studied by this perturbative method was not the Feynman kernel K ( x " , x~; T) itself, but instead the expression
Wo(x')
.
(4.1.1)
This is not a serious drawback as long as the calculation of wave functions and energy levels is concerned, but it is, of course, a more conceptual one because one is interested in the entire Feynman kernel and not in an "averaged" one. As already said, the calculation of K(T) for the Coulomb potential is not possible in the rigorous sense that the Feynman kernel can be expressed, say, in terms of elementary functions such as the harmonic oscillator. An exact expression for the Coulomb potential can only be achieved for the resolvent kernel G(E) (energy dependent Green function), the Fourier transform of K(T). Another drawback of W0 in (4.1.1) is that W0 only contains the wave functions and energy levels for even parity. Due to the general properties 1 It must be noted that Gutzwiller studied the hydrogen atom in 1967 [479] by means of path integrals. He found an exact Green function for the bound state energy levels in polar coordinates in momentum space by using a semiclassical approximation. Actually, GutzwiUer developed in these papers his periodic orbit theory, see Chap. 5.
4.1 Path Integration and Perturbation Theory
125
of the Feynman kernel, wave functions with tP(x") =  t V (  x " ) do not contribute to W0. This, however, can be circumvented by considering instead the quantity f Wl (x') = / dx"x"  Vx, K (x", x'; T) , (4.1.2) so that only states with odd parity contribute to W1. The whole picture then emerges from a proper combination of W0 and Wt. We present the improved method for a perturbative calculation of the entire kernel K ( T ) following Goovaerts and Broeckx [405]. From the result it is then obvious how to derive the two quantities W0 and W1, respectively. The general method for the timeordered perturbation expansion is quite simple. Let us assume that we have a potential W(x)  V(x) + V(x) in the path integral and suppose that W(x) is so complicated that a direct path integration is not possible. However, the path integral K (v) corresponding to V(x) is assumed to be known. We expand the integrand of the path integral containing l)(x) in a perturbation expansion about V(x). The result has a simple interpretation on the lattice: the initial kernel corresponding to V(x) propagates during the shorttime interval e unperturbed, then it interacts with V(x) in order to propagate again in another shorttime interval e unperturbed, and so on, up to the final state. One then obtains the following series expansion (see also, e.g., [65, 68,340,404408,430, 621,642,830,876], x ~ IRD)
x(t,)=x' = K(v)(x " , x ' ; T ) + ~

~..
n=l
x K (v) (xl, x'; t l ) V ( x l ) K
9. . x
+ Z
x K(V)(xt,x';tl
tn_I)V(x,)K(V)(x",x,~;T
 in)
dxj

n=l
, dtj
( v ) ( x 2 , x l ; t 2  11) x . . .
V(xn_,)K(V)(xn,x,~_l;tn
= K(V)(x"'x';T)
D dxj 
j=l
 t')V(xl)K(V)(x2,xl;t2
 tl) x ...
9.. x V(x,,_I)K (v) (x,, x,1; t ,  t , _ i ) V ( x n ) K ( V ) ( x
", x , ;t"  in) 9
(4.1.3) In the second step we have ordered the time as t ~  to < tl < t~ < ... < tn+x  t " ( T  1"  1') and paid attention to the fact that K ( t j  t j  1 ) is different from zero only if tj > t j_ 1.
126
Perturbation Theory
We consider a Ddimensional path integral with a potential V(x), x E IRD, apply the above perturbation expansion and get (we use the Euclidean path integral, see p. 34)
~x +V(x) at
(4.1.4)
x(t,)=x, Expanding the exponential yields (n = 0 term  1) oo (  h )  n
KE(X",x';T) =
Z
n!
n=O
x(t')=x"
t"
t"
n
(4.1.5) x(t')=x' Introducing the Fourier transform V(k) of the potential V(x) V(x)

1 /~ (2:r D D dk
e i k.x/h
V(k) ,
(4.1.6)
we obtain for the path integral
n=O
n!
j=l
(2~.h)D
dtj JU
D
(4.1.7) Pn(x",x';T) =
xCt")=x"/ [ 1 ftt"(2 ) ] :Dsx(t)exp  ~ , 5 2  k . ( t ) . x dt x(t,)=x, (4.1.8)
with the Ddimensional vector n
kn(t) = i Z ~ ( t  t j ) k j .
(4.1.9)
j=l
The path integrals Pn (T) are path integrals for a linear potential, and therefore can be exactly determined by solving the classical EulerLagrange equations and inserting this solution into the path integral solution for the general quadratic Lagrangian. In order to perform the nfold time integrations in the perturbation expansion, one exploits the convolution theorem of the Laplace transformation. This finally leads to the following two alternative expressions for the perturbation expansion (c > 1)
4.2 Summation of the Perturbation Series
'c
27ri Jcioo dT
esT/t~ /
127
~)Ex(t)exp
/]
 ~ ft,
~k2 + V(x) dt
x(t')=x, {DO
L ,0/o
:, (2,~h)D
:, (2,~h)~
n.0
D (2.h)~'
exp (  ~ x ' . E jn= : k ~
•
i , , k0) fix 9

(4.1.10)
Is + (k0~/2m)]... Is + (k0 + . . . + k.)2/2m] oo
=
E(_l)n/~" ,.=o
n
dk0
j_I~I/R dkj V(kjkj_:)
:, (2,~h)~ _
•
D (2~h)~'
exp [~(x' 9kn  x" 9k0)] (~ + k0~/2r~)... (~ + k~./2m)
(4.1.11)
This method therefore allows one to perform the path integration as explicitly as possible. In the final expression path integrations no longer occur. However, one is left with convolutions in Fourier space together with an infinite summation. It is obvious that such a complicated formula allows the necessary final manipulations only for particular problems. Examples can be found in [401,404409]. 4.2 Summation
of the Perturbation
S e r i e s f o r 6 a n d $ '  P o t e n t i a l s
In Sect. 3.3 we discussed some basic path integral solutions. However, there are several potential and boundary problems which do not fall into these classes and are exactly solvable quantum mechanical problems nevertheless. The (ifunction potential and potential problems with a (ifunction perturbation belong to these kinds of problems. Based on a functional analytic approach, point interactions as solvable quantum models have been thoroughly discussed by Albeverio et al. [17]. In this section we discuss all these nonGaussian path integral problems which can be interpreted as path integrals with point interactions, respectively boundary conditions. As it turns out, the corresponding Green functions can be written as a quotient of two determinants with the unperturbed Green functions taken at the perturbation points as entries [17,439], cf., the section about point interactions in the table of path integrals. However, this simple feature does not hold for multiple point interactions in the path integral representation for the onedimensional Dirac particle. The former simplification does not work in general, and the corresponding Green functions are matrices with determinants as entries which are in turn also determinants, and so on, cf. [448].
128
Perturbation Theory
4.2.1 O n e  D i m e n s i o n a l Point Interaction. In order to incorporate a onedimensional point interaction, i.e., a onedimensional Jfunction perturbation we consider in (4.1.3) the following potential in the path integral [430] w(x)
= v(z)
 ~(x
 a)
.
(4.2.1)
The path integral for this potential problem has the form
KCx"'x';T) =
t
I.,
z2_ WCz) dt
C4.2.2)
,(t,)=,' Let us assume that the path integral for the potential V(x) is known, i.e., the path integral
K(V)(x", xl ; T)
.
:=
.(t)exp
,
(4.2.3)
~(t,)==, and also the (energy dependent) Green function
G(V)(x",x';E) = ~ i f o ~ dT eiETIh K(V)(x",x';T) K(v)(x"'x';T) =/n~ ~ri
(4.2.4)
iET/hG(V)(x,,,x,;E )
Introducing the Green function G(E) of the perturbed system similarly to (4.2.4) we are able to sum exactly the perturbation expansion, due to the convolution theorem of Fourier transformation, yielding
G(6)(z',x';E) = G(V)(x",x';E)
_
e(V)(x"'a;E)G(V)(a'x';E) G(V)(a,a;E) 1/7
(4.2.5)

Here it is assumed that G(V)(a, a; E) actually exists and the energy levels E (~) of the perturbed problem W(x) are therefore determined in a unique way by the equation aCV)(a, a; E (~)) = 1 . (4.2.6) 7 This is in general an implicit (transcendental) equation. The corresponding wave functions are given by ~P~6)(x) = [ lim
E}~ )  E
)
1/7_G(V)(a,a;E )
]1/2
GCV)(z,a;E} ~)) .
(4.2.7)
4.2 Summation of the Perturbation Series
129
4.2.2 P o i n t Interaction for the Dirac Particle. We consider a point interaction in the path integral for the onedimensional Dirac particle. From the theory of [508] we have for the onedimensional Dirac particle
K(v)(x",x';T) =
/
Dv(t)exp
 h
W(x)dt
,
(4.2.8)
~(t')=x, with W a matrixvalued potential, and the Green function G(V)(E) in its matrix representation is defined as G(v)(x " , x ' ; E ) =
( 'q(v)/~''' E) "'11 ~ ,x';
~(v), ,, E)) (;12 [x ,x';
m(v)r~'" E) n(v)r" E) "  ' 2 1 I,~ ,X/; "  ' 2 2 ~ x ,,~1;
(4.2.9)
We first consider a Jfunction perturbation in the electron (= "+") component, i.e., V =  4 ( 0 1 0)J(x 0  a). We obtain by inserting it into the path integral and summing the perturbation expansion 1
G(6+) (x ", x'; E) = G(v)(x '', x'; E) + 
( G~V)(a,x';E)G~V)(x",a;E) x G~V)(a,x,;E)G~)(xU, a;E)
a; E )
r:'_(v) ~.,. 1~r:(v)(x,t,a;E) "~11 ( a , .~ , ~~]~'12 r.=,(v) ~,., ,''12 ~,~(y) (x", a; E) ] (4.2.10) 9~21 (a, .~
Similarly for the positron (= "  " ) component, i.e., V = (4m2/3c2/h 2) (o ~)J(x  a) (the constants have been chosen for convenience) G(~) (x '', x'; E) = G(v)(x ", x'; E) 
1
h2 / 4m2c2 /~ k G~Y) (a, a; E) f n(v)rx'" ~x~(Y)(z", a; E) G(Y)la 12 ~ ' x l ; E~G(V)Ix ] 22 I, u a ; E ) ) 1 x "~12 (v) ~ , ,.,  " / "  ' 2(y) 22 ~ ,a;E) (4.2.11) \G22 (a,x,E)G21 (x",a;E) G(Y)[a22, ,x';E~G(V)Ix
The investigation of the two problems shows that in the nonrelativistic limit the path integral incorporation of the J and J'function perturbations, respectively, emerge. The former case has already been discussed in Sect. 4.2.1, the latter we state in the next section. 4.2.3 O n e  D i m e n s i o n a l $~Function. The case of the incorporation of a J'function perturbation is as it stands not well defined due to the ultraviolet divergence. In order to make the problem well defined one regularizes it by considering a Jfunction perturbation for the onedimensional Dirac particle (see above) and its corresponding path integral formulation. The nonrelativistic limit of the point interaction in the positron component is
130
Perturbation Theory
due to the particular feature of the emerging boundary condition called a ~f'function perturbation. Performing in (4.2.11) this limit yields the following result for the corresponding Green function [446,448]
i ~oo dT ei ET/h
G (~') (z", z'; E) = ~
=(t")==" x
/
[ i t"
"Dx(t)exp ~
(2x2V(x) W~'(xa))dt
]
=(t')==' ~(v) r
.
(v)
,.
= G(v)(x,,,x,;E)_ "',z, ," ,a E)G~:,,(a,x ,E) 1/t3+~(v)=,,(a,a;E) ,
(4.2.12)
G!V)(a'a;E)= ~O2k, OxOyG(V)(x'y;E) 2m~~(xY)) 3:=y=a
(4.2.13)
Actually, a point interaction in one dimension has a fourparameter family of selfadjoint extensions [17, 73, 144, 145]. These four parameters can be successively incorporated into the path integral, cf. [448]. 4.2.4 B o u n d a r y V a l u e P r o b l e m s . Actually, every point interaction in the path integral corresponds to a particular boundary condition. However, in the following we want to discuss two specific kinds of boundary conditions, i.e., Dirichlet and Neumann boundary conditions, respectively. We present the incorporation of (Dirichlet or Neumann) boundary condition at one point a E IR, where the motion can take place on the right or left hand side of a, i.e., x > a or x < a, respectively. To take into account a second boundary condition b E lR we just repeat the procedure, and the motion is restricted to a < z < b. This approach turns out to be extremely useful for potentials which are restricted to z _= r > 0, i.e., potential problems which are defined and solvable in IR are interpreted as radial potentials. This restriction can alter the properties of a potential completely, for instance, from a pure scattering potential to a confinement potential [439]. If the correct boundary condition is not taken into account, the results can be very misleading [275].
~.2.~.1 Dirichlet Boundary Conditions. In
(4.2.5) we consider the limit 7 +  c o which has the effect that an impenetrable wall appears at x = a [439]. We set lirrb~_oo G(~)(E) =_G(==a)(E),i.e., we obtain
G(==a)(z ", z'; E)
x(t")=x" =hfo ~dT eiET/~ / 79(D)(==a) z(t)exp x(t')=x'
] [h ft,t" ( 2 z 2 _
V(z))dt
4.2 Summation of the Perturbation Series
131
= G(v)(x ", x~; E)  G(v)(x"' a; E)G(V)(a, x~; E) G(V)(a, a; E)
(4.2.14)
Bound states are determined by the equation
G(V)(a, a; En) = 0 .
(4.2.15)
Again, the wave functions are given by the residua of G(E) at E = E,~, i.e.,
[ G(V)(x'a;E)] E,E,~  ( E n  E ) G ( V ) (a, a; E))
~,(x) = lim
(4.2.16)
4.2.4.2 Neumann Boundary Conditions. In the case of Neumann boundary conditions one can start by considering a 6~function perturbation in the path integral and by making the strength infinitely repulsive [17] after summing up the corresponding perturbation expansion. We therefore obtain a path integral formulation for Neumann boundary conditions at x = a (superscripts (D, N) denote in the following Dirichlet, respectively Neumann boundary conditions, for the Green function) [446] G (N) (x", x'; E) x(t")=x 1'
. = . ~ o ~ d T e iET/h x(tl)=~ , ~(y), ,, E)G(~)(a ' E) = G(Y)(x.,x~;E ) _ u,~, (x ,a; , x'; , O(v) , a", E) ,x,x,,(a, G(V)tzy, a, , a; E) =
(4.2.17)
G(V)(x, y; E)  ~6(x  y) (4.2.18)
Bound states are determined by the equation ~(v) ,x~x. ,[ a ,
a;
E,)

0
(4.2.19)
.
The wave functions are given by the residua of G(E) at E = E,~, i.e., ~,, (x)
lim [(E~ E  ~ E .
G'v)(x'a;E)]  E ) ~ ! y ) ( a ,
a;
E)
(4.2.20)
132
Perturbation Theory
4.3 Partition Functions and Effective Potentials Let us consider 2 the path integral for the onedimensional density matrix (see (2.2.11))
~(o)=~'
An interesting object in statistical mechanics is the partition function Z := i ~ p(x, x; 13)dx = Tre  z s
(4.3.2)
With (4.3.1) we obtain for the partition function
.(,,: fo.. 0 semiclassical expansion of the Ddimensional Feynman kernel
the following
K(x",x';T) (27rih)D/2 1 ~ ldet ( • Here several remarks are in order:
a2P~ ,b)] 1/2 ox,,oo .(l+O(h))
.
(5.2.10)
146
Semiclassical Theory
i) The orbit formula (5.2.10) gives the leading term of the Feynman kernel in the semiclassical limit h ~ 0 as a sum over all classical trajectories, also called orbits, which connect x' and x" in time T. ii) In deriving this formula from the path integral (5.2.2), the Feynman paths x(t) have been expanded about the stationary points, i.e., x(t) has been decomposed into x(t) = xT(t ) + q(t) with q(0) = q(T) = 0: q(T)=0
K(x",x';T) = ~~ /
(5.2.11)
2)q(t)eiR[x~(t)+q(t)]/h
J
"t q(0)=0
Then the action R[x~ + q] has been expanded about the given classical path to second order in the quantum fluctuation q(t) leading to the path integral for a Ddimensional harmonic oscillator with time dependent frequencies whose path integral solution is known, see Sect. 3.2. The result is 1
(1
e iR~(x'',x';T)/hiru~/2
K ( x " , x ' ; T )  (27rib)D~2 ~.y
idetJ.~(T)ll/2
9 \ +O(h))
(5.2.12) where the matrix valued function J~(T) turns out to have a geometric interpretation as the socalled Jacobi field, see e.g. [828], of the classical trajectory emerging from the initial point x'. Furthermore, the number ~7, which enters the above derivation as the number of negative eigenvalues of the matrix J r ( T ) , is known as the Morse index of the trajectory and has the geometric meaning of the number of conjugate points on the trajectory x.r(t), 0 < t < T, conjugate to x', counted with multiplicities. In the following we shall present only a short discussion of this aspect of classical mechanics which deals with the calculus of variations in the large of classical trajectories. iii) Consider the family x(p';t) of classical trajectories which start at x' at time t = 0, x(p';0) = x', parametrized by their initial momenta p' = mx(p';0). Thus x(p';t) are trajectories spreading out from x' in different directions with different velocities. One defines the D • D matrix J ( p ' ; t ) by Jkt(p';t) := 0xt(p';t)
Op'k
(5.2.13)
Since x(p'; 0) = x' for all momenta p', one obtains that J(p'; 0) = 0, and furthermore Jkl(P'; 0) = 10p~ = ltikl . (5.2.14) m tgp~ m The matrix J (p'; t) now describes the deviation of a trajectory x (p' + e; t) from x(p~; t) at t > 0 to first order in e, xk(p' + e; t) = Xk(p';t) + Jkt(p';t)et + O(e 2) 9
(5.2.15)
5.2 Semiclassical Expansion of the Feynman Path Integral
147
All trajectories x(p'; t) are solutions of the equation of motion +
OV
= o,
(5.2.16)
which yields after taking the derivative with respect to p~ the Jacobi
differential equation
+
o 02V(x(p,;
ox ox, tllykn(P';t) = 0
(5.2.17)
n1
Thus the Jacobi field J(p'; t) satisfies the differential equation
m'J(p';t)+V
(5.2.18)
with initial conditions J(p'; 0) = 0 and J(p'; O) = n/m. The uniqueness of the solutions of the Jacobi equation (5.2.18) implies that the matrix valued function J~(T), whose determinant enters the semiclassical formula (5.2.12), can be identified as the Jacobi field along the classical trajectory x~, J~(T) = J(p';T). The geometric meaning of the Jacobi field can be visualized by a change of the point of view: presently we consider the problem of investigating all solutions x~ (t) of the classical equations of motion with fixed boundary conditions at t = 0 and t = T. Now we view a given trajectory xu (t) with initial conditions x~ (0) = x' and mx~ (0)  p' as a function of the endpoint of the interval [0, T], i.e., x~(T). The determinant of the Jacobi field J ( p ' ; T ) of this trajectory vanishes at tc if there exists another trajectory x(p' + e; t) with initial conditions x(p'+e; 0) = x' and mx(p'+e; 0) = p'+e that intersects x~ (T) at T = to, i.e., x(p' + e;tc) = xc = x~(tc). This intersection takes place at the focal point xc which is also said to be conjugate to x' = x~(0). For small times T such that on x~(t) there occur no conjugate points in the interval 0 < t < T, all eigenvalues of J(p'; T) are positive. This result derives from the fact that then the trajectory 7 minimizes the action functional (5.2.8) in Hamilton's principle. At a conjugate point xr at least one eigenvalue vanishes, and the multiplicity of this zero mode is given by the dimension of that subspace of momentum space for which x(p' + e; t~) = xc. Increasing T then turns these eigenvalues of J(p'; T) negative. Thus, for arbitrary T the number v~ of negative eigenvalues of J~(T) = J ( p ' ; T ) , the Jacobi field along the classical trajectory x~(t), is the number of conjugate points on x~(t) for 0 < t < T counted with multiplicities. The number v~ of conjugate points can be identified with the socalled Morse index of the trajectory x~(t) as a result of the Morse index theorem [697,713, 828]. iv) The final step in the derivation of (5.2.10) consists in expressing the determinant of the Jacobi field J.~ (T) evaluated at the final time T by Hamilton's principal function P~, see (5.2.9). It is well known that P~(x", x'; T)
148
Semiclassical Theory is the generating function of the canonical transformation, commonly labelled as type 1, which corresponds to the classical time evolution in phase space backwards in time, i.e., (p", x " ) ~ 9 (p',x') with x" and x' as independent variables, and the final and initial momenta p" and p', respectively, as dependent variables. The latter are given by the relations p " = Vx,,R 7 ,
p'= V,,,R 7 .
(5.2.19)
The corresponding classical trajectory x~ (t) connecting x' and x" in time T has energy E.~ which can be calculated from R 7 as follows E~ 
(5.2.20)
aT
With 1
we derive detJ.~(T)= [det(
02R7
~]i,
a ,,oOx,b ] j
(5.2.21)
and thus (5.2.12) is identical to the semiclassical expansion (5.2.10). v) It has become a widespread though inaccurate custom to call the determinant which occurs through its square root in the semiclassical formula (5.2.10) for the Feynman kernel the Van Vleck determinant on the basis of Van Vleck's work of 1928 [906]. An improved designation of this determinant is sometimes given as the Van VleckPauliMorette determinant [613, Sect. 4.3]. It even happens that the semiclassical formula itself is unduly called Van Vleck's formula. Recently, convincing evidence has been provided [193] that the semiclassical formula (however, for small 3 times T only and without the summation over the classical paths 7 and the corrections coming from the Morse index) is due to Pauli [765], and that the determinant which occurs in Pauli's formula is, up to a sign factor, the same as that due to Morette [710], and Van Hove [905]. Thus in the following we call the semiclassical expansion (5.2.10) Pauli's formula and the determinant the MoretteVan Hove determinant. See also the historical discussion in Chap. 1, p. 11. vi) Notice that Pauli's semiclassical formula (5.2.10) is exact for systems whose Lagrangian is at most quadratic in x like the free particle (2.1.59) and the harmonic oscillator (3.2.24). 3 Here small T means T < t m with tm := min~t2, where t~ denotes the time at which there exists on a given trajectory x~(t) the first conjugate point conjugate to zT(0) = ~'.
5.3 Semiclassical Expansion of the Green Function
149
vii) The singularities of the MoretteVan Hove determinant which arise from the zero modes of the Jacobi field at conjugate points were investigated for the first time by Choquard [192] for conservative systems with nonsingular, confining potentials. This is a class of systems which allows an infinity of trajectories passing through x' at time t' = 0 and x" at time T. It was shown in [192] that the manifold of conjugate points is given by cgx'k/cgE.y = [cg~R~/Ox'k OT] 1 = O. The first person to go beyond the conjugate points was Gutzwiller in 1967 [479] who made the formula (5.2.10) as the starting point for the derivation of his trace formula, see Sect. 5.4. For a rigorous mathematical derivation of the semiclassical formula (5.2.10), see [477]. 5.3 Semlclassical Expansion of the Green Function
The Green function G is defined as the Fourier transform of the Feynman kernel, see (2.1.25) G(x",x'; E ) : = ~i ~oo dT ei(E+i e)T/ti K (x", x'; T)
(5.3.1)
A formal way to derive the semiclassical asymptotics of the Green function is to insert into (5.3.1) the asymptotic expansion (5.2.10) for K, then to interchange the summation over classical paths with the time integration, and finally to evaluate the remaining Fourier transform by the method of stationary phase. However, there one encounters a serious problem. It is well known that the two limits h ~ 0 and T + ~ do not commute. On the other hand, one requires in (5.3.1) the kernel for T + cx~,and thus the semiclassical formula for K, which holds for h + 0, cannot be used directly. Since both G and K are distributions, one should regularize them with suitable test functions. Here we shall ignore this problem, but shall come back to it in Sect. 5.4. Inserting the semiclassical expansion (5.2.10) for K into (5.3.1) leads to G ( x " , x ' ; E )  hi (2~ri lh)D/2 ~
fo~dTI det
(  Ox'/aOx) 112
Evaluating the time integration by the method of stationary phase, we obtain 0
OT
[ P ~ ( x " , x ' ; T ) + E T ] =0
(5.3.3)
and with aP~/aT = E~ the following condition for the stationary points in T for a given energy E
150
Semiclassical Theory (5.3.4)
E~ = E~(x",x';T) = E .
The solutions to this condition are the travelling times T~ = T~ (E) of classical trajectories x~(t) which go from x' to x" in time T~ with energy E. While in the semiclassical formula for K the energy of the classical trajectories was not fixed, the leading semiclassical contribution to the Green function G comes just from those classical trajectories which possess fixed energy E. Evaluating the integral in (5.3.2) at a given stationary point T = T~ > 0, we obtain the following phase
(5.3.5)
P~(x", x'; T.~) + ET.~ = Su(x", x'; E) : = / ~ p . dx ,
which is just the classical action of the given trajectory 7 with energy E. The final result for the semiclassical expansion of the Green function reads G(x", x'; E) = G(x", x'; E) + • E
where
i
1
li (2rih)(D1)/2
~/D~(x",x'; E)exp [ ~ S ~ ( x " , x ' ; E )  i 2 / ~ ] . (1 + O(h))
I D~(x", x'; E) := det
02S OxHaOxtb r S.y
ax' aOE
(5:3.6)
02S 02 S.y
,
(5.3.7)
OE2
and (~ denotes the contribution to the Green function from the stationary point at T = 0. Here the index #u counts the number of points on x~ conjugate to x' in energy E rather than in time and is defined as ]'v7 , u~+l ,
D2R~>0 , D~R~ < 0 ,
(5.3.8)
with D~ P~ denoting the second derivative o f / ~ :
D~P~ := OT2I~(x",x';T) 92 T=T.y
(5.3.9)
In the above derivation we have assumed that the stationary points are nondegenerate, i.e., D~R~ ~ 0 which states that x" must not be conjugate to x' in energy on any of the trajectories x~.
5.4 The Gutzwiller Trace Formula
151
5.4 T h e G u t z w i l l e r Trace Formula In seeking a substitute for the EBKquantization condition (5.1.2) in the case of strongly chaotic systems, Gutzwiller [479] concentrated on the spectral
density 4 oo
d(E) := y ~ a ( E 
En) 9
(5.4.1)
n0
Using the spectral representation (2.1.50) 5 E) = n0
~. (x")~,; (x') = n=o P
E7E
+i~r~~,(x")~,*(x')a(E~,=0  E) (5.4.2)
we observe that the spectral density d(E) can be obtained from the trace of the Green function oo
1
_ dxG(x,x;E) Tr(H  E  ie)l = ~'~ E,  E  ie = f,,~D
(5.4.3)
nO
as follows 1~ f H(E) = ~ J~D d x G ( x , x; E)
(5.4.4)
In general, the resolvent of I=I is not of trace class which manifests itself as a divergence of the infinite sum in (5.4.3).6 This problem can be overcome, however, if one considers a smeared level density with a suitable test function; see remark vi) below. To obtain the Gutzwiller trace formula, we insert the semiclassical expansion (5.3.6) for G into (5.4.4) 7
d(E) = d(E)  2D { (2~rili) (D+1)/21 ~
x exp g & ( x , x ; E )  i T .
f=jD,(x,x;E)o
, 9
dx
+O(h))
(5.4.5)
where the sum over 7 runs over all closed classical trajectories x~ starting out at x and returning to this point after a time T~ = T~(E) > 0. 4 Here we assume that the quantum system whose classical limit is described by the classical Lagrangian (5.2.1) has only a discrete energy spectrum E0 _< E1 _< . . . . 5 Here P denotes the principal value. ~ E.g., for twodimensional Euclidean billiards it follows from Weyl's law [49, 919] that En = O(n) for n + ~ . 7 d(E) is the contribution derived from the term O in (5.3.6); see remark v) below.
152
Semiclassical Theory
The method of stationary phase applied to the integral in (5.4.5) leads to the condition [Vx,,S~ (x", x'; E) + Vx, S~(Z', x'; E)]
xll~xl,~_ x
=0
(5.4.6)
for x E IR~ From the definition of the classical action S~ as a Legendre transformation of R~, see (5.3.5), one derives VxSw(x",x';E)=p",
Vx, S ~ ( x " , x ' ; E ) =  p ' ,
~~Sw(x O ,, ,x , ; E)
=
Tw,
(5.4.7) and the condition (5.4.6) yields p"  p' = 0, i.e., the closed trajectories x~ must have identical initial momentum p' and final momentum p". Thus the condition of stationary phase applied to (5.4.5) picks out all points x in configuration space that lie on some closed orbit xw with T~ > 0 and with the additional property that initial and final momenta are equal. That is the stationary points are those orbits which are closed in phase space, i.e., the periodic orbits 3" of period Tw > 0. These stationary points can never be isolated since a periodic orbit 3' itself is a onedimensional manifold of stationary points, along which the classical action Sw(x, x; E) is constant, i.e, for x on the periodic orbit 3' Sw(x, x; E) = f~ p . dx =: S~(E) .
(5.4.8)
Here we shall not give the somewhat subtle calculation of the integral (5.4.5), but rather refer the reader to the literature [109, 111,479,480,483]. We then obtain the Gutzwiller trace formula d(E) = d ( E )
1 ~ + ~~
T~r (1 ~rk~ ) ( ) = i det(Mk _ 11)11/2cos ~ k S ~ p  ~ ~u~p 9 l + O ( h )
+ O(h ~176.
(5.4.9)
Here several remarks are in order i) The sums in (5.4.9) run over all primitive periodic orbits 39 with period T~p, and their kfold repetitions, k E lN.s ii) The matrix M.~ is the socalled monodromy matrix and is given by a linearization of the Poincard recurrence map P.y~. To simplify the discussion, we mention only the case of systems with two degrees of freedom. Then M r is a 2 x 2 matrix with two eigenvalues Aw and A~ 1. Furthermore, since det M.y = 1, s A single traversal of the set of points on a periodic orbit is called the primitive periodic orbit "/p corresponding to %
5.4 The Gutzwiller Trace Formula
153
det(M~  11) = 2  TrM~ = 2  A~  A~ 1 ,
(5.4.10)
which implies that A~+I  21( T r M ~ t ~f(Tr M~) 2,  4 )
.
(5.4.11)
We distinguish four cases: a) T r M ~ > 2 : In this case one obtains two real eigenvalues A~ 1  e +u~, u~ > 0, and 3' is unstable. (A nearby trajectory locally separates away from 3' at a rate e x~T~, where ~ = u~/T7 is the Lyapunov exponent of 7.) Such a periodic orbit is called hyperbolic, and I d e t ( M ~  n)[ 1/2
2sinh uz~
(5.4.12)
b) Tr M r <  2 : Again one obtains two real eigenvalues, but now negative ones, A~ 1 = e+U~, which again means that 7 is unstable. In this case the periodic orbit is called inverse hyperbolic, and I det(M~  li)l 1/2 = 2cosh ~
(5.4.13)
c) I Tr MTI < 2 : Now both eigenvalues of M7 are complex, A~ 1 = e + i.~, v7 E (0, zr). 7 is a locally stable orbit (in the linear approximation of the dynamics) and neighbouring trajectories wind around it. In this case 7 is called elliptic, and ] d e t ( M ~  11)11/2 2sin ~2
(5.4.14)
d) I TrMTI = 2: This case implies that both eigenvalues are either A~ 1 = +1 or A~ 1 =  1 . Such a periodic orbit is called parabolic. Once A.r = 1, one observes that the trace formula becomes inapplicable since then det(M, r  11) = 0 and the respective term in the trace formula diverges. Otherwise I det(M~  11)11/2 = 2. iii) In the derivation of the Gutzwiller trace formula (5.4.9) we have assumed that all periodic points x E IRD are contained in smooth connected onedimensional manifolds, i.e., all periodic orbits are single isolated periodic orbits. iv) The index/]7 is called the Maslov index and denotes/~.~ plus the number of negative eigenvalues of the second variation of the classical action $7 evaluated in local coordinates along the periodic orbit. v) The first term in the trace formula d(E), which is due to the singularity of the Feynman kernel at T = 0 and is proportional to vol(I2E), corresponds to the ThomasFermi approximation in the general case, and to the Weyl term [49,919] in the integrated level density for billiards. Here vol(I2E) denotes the volume of the energy hypersurface I2E which is defined by DE = {(p, x ) l g ( p , x) = E} of energy E in phase space
154
Semiclassical Theory vol(/2E)
=/~vdp/~v
d x ~ ( E  H(p,x))
v(~) O h
=
r2/al
R = 0"2z 2, r > 0
2m ~ e~r
+
h~ {g:
g~/v'az.
ffl

,
h = 2~/~R = aoz 2 + b o z + c o
xEIR
R=I,
h ~ f z ( z  l) + b o o  z) + h, 2m
R(z)
+ aV(z),
li2
f/4 ) \
z = 1(1 + tanhx)
z ~ (o, 1)
cosh 2 x
I
R=z,
\
/
r>O
z = tanh ~ r
h2 [ f3/4 h0+hiH3/4 2m t h l t 1 + cosh2~ + sinh 2 r
R=lz, r>0 z = 1/cosh 2 r
h2 ' 2m h 0 + h l + l
R   z 2, r > O z =
1  e 2r
n = 4z(1  ~), ~ = ( 0 , . ) z = ~(1  cos ~) &
k
h2 ( h i  ~4
2m
f H3/4 h_l F3/__4 cosh2r + sinh 2r ]
~
+ f + l + 4sinh   2rF
h 2 ( h o F h, F 3/4
Sm ~ ~i~Tj~
)
f
)
cothr
hlq'3/4
+ cos~(~/2)
(f + 1)g
)
160
Table of Path Integrals
r
I
b~
~9
g.
~9
+ ,
O(T)
6.1.1.1 PrePoint Formulation [69,76,173,228235,264,265,314,368,416,494, 565,568,636,637,690,761,797,899,914,915] after DeWitt (DW):
[
q(tH)=q"
t"
Rh2 ,~dt] "
q(tl)=q ~
(6.1.1)
6.1.1.2 Symmetric Rule (SR) [26,76,264,266,690,888,894,939]:
q(t')=q ~
(6.1.2) 6.1.1.3 MidPoint Formulation (MP) [198,217,260,346,358,359,380,389,392, 464,494,565,568,572,593,647,648,645,665,676,698,736,811,815,888,914,915]: q(/.)__qU
= [g(q,)g(q.)]l/4
i
7)MPq(t)x/g
q(t')=q'
x exp ~
,
)7
gab(q)qaqb  V(q)  AVMp(q) dt
(6.1.3)
6.1.1.4 ProductForm Formulation [422,447,470] (hachcb = gab):
q(t')=q'
6.1.1.5 Vielbein Formulation
[611613,616] (d~ ~ = e~(q) dq", ~ ;
(6.1.4) = ~a;. ~qj.
e~;g,u Aq~*Ajq v 12 + ~;.,~ Aq~ Ajq v Alq~./6): =
lim t2~~eh) N~
j ~ dAx~exp ~ _
~,2c
'
(6.1.5)
AVMp = AYweyl,AVsrt, AVpF as in (2.8.19), (2.8.20) and (2.8.40), respectively, and R is the scalar curvature.
6.1 General Formulae
163
6.1.2 PhaseSpace Formulation, Hamiltonian Path Integral. [26,135,
245,248,340,375,380,394,564,621,634,635,677,685,701,703,704,732,772,788,828, 862,894] ) D_ 89
(  2V/~ z< ) . (6.2.37)
The following recurrence relation holds for the DdimensionM harmonic oscillator kernelK (D) (T) (v = x'. x")
K (D)(x", x'; T) 
1 0 K(D_2)(x,,,x,;T) 2~r cgv
(6.2.38)
"
The Green function is given by (~ = ~(]xl' + x"] + ]x"  x'D, 17 = ~([xl, + x" I  I x "  x ' l ) , D = 1,3,5...) i ~ aT eiET/h hf0
~(r /
[
r ] im w2x ~) dt /)x(t)exp 2hft, (x2
x(t')=x, D1
DI 2
(6.2.39)
180
Table of Path Integrals
6. 2.1.8 Repelling Harmonic Oscillator.
[146,281,359,485,664,753]
z(t")=x" 7?x(t) exp
~
(/:2 + w2x2)dt

2ih['
~in~TJ J '
(6.2.40)
1 h V/~fr t dEexp (  h T E + ~ wr Eh ) = 87r2 
~+i E/hw
+ F ( ~ + iE'~[2E(1 , 2t~) I  89
( ~
5.2.1.9 Forced Harmonic Oscillator.
"/)
" ~ ( 1 ) ( ~ "~
x(T)=x" K(x", x'; T) =
/ /)x(t)exp l hR[X(t)]} x(O)=x,
_ (2rrih,D/2 1 ~
I det (
02/Lr
I
Oz""Oz 'b J ,
•
.(l+O((h))
. (6.2.143)
Here the sum over 3' runs over all classical paths z~(t) satisfying the endpoint conditions x.~(0) = x', x~ (T) = x". Since the time T is fixed, but not the energy E of the classical paths, there usually will exist several solutions to Hamilton's principle. R~ denotes the classical action evaluated along an actual path x~(t) of the system R~ = Rv(x" , x'; T) := R[x~(t)] ,
(6.2.144)
where The number u~ is known as the Morse index of the trajectory and has the geometric meaning of the number of conjugate points on the trajectory x~ (t), 0 _< t 0, satisfying h(p) = O(p2~), Ipl ~ c~,5 > O. (#a is the socalled "entropy barrier", see [37, 109, 866, 868, 869].) 17p denotes the geometrical length of 7p, and X7~ is a phase factor depending on the Maslov index 7p and the boundary conditions, e.g., Dirichlet or Neumann boundary conditions. C is the constant in Weyl's improved law for billiards, see e.g., [45,49]. g(u) = (1/Ir) f ~ h(p) cos(up) dp.
6.2.4.3 Trace Formula for the HeatKernel on the Sphere  Orbifold SpaceTime IR x S ( 2 ) / F . [174] KS(2)/r(T)

Irl (rr) 3/2P
da
sin' ~
or+ 2 {7}

(6.2.185) Here S( 2 ) / F is the fundamental domain of F (an elliptic triangle), q is the generic order of the rotation such that for each primitive 7 E F we have 7q = 11, where F is a finite subgroup of 0(3) acting with fixed points, and nq is the number of conjugate qfold axes. The summation over all primitive conjugacy classes of elements 7 E F is denoted by ~{7}" The quantity Irl is defined via Irl = 4rr and describes the order of F.
fs(~)/rds
6.2.4.4 Trace Formula on C P N Manifolds. [355] (Q = }']~ n , with n~ the occupation numbers, c~ are the matrix elements of the Hamiltonian _H = a t diag(c~,..., CN+l)a_ of the coherent states a_ = (_al,...,_aN+~) with the commutation relations [_a~,_a~] = 6~Z 11, [_aa, _a~] = [_a~,_a~] = 0, V,,Z) N+I
e i Qc,,/~
Z(fl) "" E Ho~*/~ /'1 a1 ~
e i`c,c~>/~ "~
)
(6.2.186)
6.2.4.5 Trace Formula for Zoll Surfaces. [825] oo
E 5 ( p  v/E~n) = E
ak(p)ei~n'k/2e2rikP
(6.2.187)
n=0 kE~ is the Maslov index, the ddimensional Zoll surfaces (= SC2.manifolds) are manifolds all of whose geodesics are closed,and ak (p) = vol(SC2,r)pa 1/(4~r)a/2 r(d/2 + 1) + O(pa2), as p ~ +oo, and c~k(p) = O(p~176as p + oo.
6.2 The General Quadratic Lagrangian
215
6.2.4.6 Periodic Orbit Formula for Integrable Systems. [92]
ImM"MI e2rimM'MU1/D88 DU(D1)/2DIMI(D1)/2 I d et 102m'M M#O
Z J ( U  U ( m ) ) = 1+ Z m
i
m=mM
(6.2.188) Here U is the total number of states below energy E, U(m) denotes the scaled levels in the regular spectrum labelled by D quantum numbers ( m l , . . . , m D ) =: m, M is a Ddimensional lattice of integers, and /J = (#1,...,/~D1) are D  1 curvilinear coordinates on the contours U(rn) = 1 in m space. The points mM := m(/AM) are defined by the points/AM where M. Om/al~a = 0, 1 < a < O  1, M = M/IMI.
6.2..t. 7 Periodic Orbit Formula for Perturbed Quadratic Lagrangians. [12] ~adxr
/ Dx(t)exp f0 xtAx2 x(0)=x (ar0) 1 oo ( i ~r ) "~ ~ E Z exp ~kRTp i~ku.y,
V(x)dt
7p k1
dt
Jo
r
dr ,
(6.2.189)
with the quantity Cp = ~
1
d1 H 4c~
T
k=l
n
i1/2
(6.2.190)
Here A is a d • d strictly positive matrix and the potential V is such that the energy E is the only conserved quantity of the system, with 7p the family of all primitive periodic orbits of period T/n, v.yp a Maslov index for 7p of the closed orbit 7 T, ~k the stability angles, DT : det(O2T/Ox~Oxb), where XT,L denote the transversal (longitudinal) coordinates with respect to the periodic orbit, and r E C~~ d) is a suitable test function.
6.2.4.8 The Selberg Trace Formula. [489,844,866,910] ~'~h(p,)= n
~ fo ~
dpptanhTrph(p) + E
oo l~g(kl~) Z sinh~
{~}~Pk=l
(6.2.191)
Here units h = 2m = 1 are used, g(u) = ( l / r ) f o h(p)cos(up) dp, E = p2 + 1/4. {7} E T' denotes the summation over all primitive conjugacy classes in the Fuchsian'group F. l.r denotes the length of a closed geodesic corresponding to a closed 7 9 F. h(p) is an even test function with h(p) = O(p2'), as IPl ~ oo, and is analytic in the strip I~(p)l < 1/2 + of,(f > 0..4 is the area of the fundamental domain corresponding to F.
216
Table of Path Integrals
6.2.4.9 The Selberg Trace Formula for Compact Symmetric Space Forms of Rank One. [363]
j>0
[c(ir)[2
{~} (6.2.192)
1
_
rC~2~)rCir + m~12)rCi,t2 + rn~/4 + m~/21 F(ma + mo)F(ir)F(ir/2 + m~/4)
c(ir)
1 =
+
'
(6.2.193) (6.2.194 /
1 
c(z) is the HarishChandra cfunction. Further, for any ~, ~ stands for the character of A = AeAp defined by ~ ( h ) = e a(l~ and e~(h) is, for h e A, equal to the sign of 1Ia~+ (1  1/~a(h)), r
being the set of roots
of (9C, aC), i.e., those that are real on a. C(h) is a positive function on A, cf. Sect. 6.10.9, and [363]. The test function h(r) must fulfill the requirements: h(r) is holomorphic in the strip I~(r)l _< t~0 + c, e > 0, h(r) has to decrease faster than Ir12 for r + +oo, g(u) = 7r1 f o h(r) cosQrr)dp. See also 6.7.14 and 6.10.9.
6.2.4.10 The Selberg Trace Formula on DDimensional Hyperbolic Space. [156,363,364,909] h(p.) = 2v n~O
fo ~ dph(p)~o(p)
+ 2Z
{'r}
l~a(l~) K1)I ' N(D1)/2ldet(ll_S_l (6.2.1951
~(P)
r~(D) r(ip + O_~)I~ (2~)2
r(ip)
(6.2.1961
!
The HarishChandra function in this case is relatively simple and gives a polynomial of degree P  ~ in p~ if D is odd, and a polynomial of degree 0.2 in p2 times ptanh~rp, if D is even. The matrices K E O(D  11 and S 2 denote a rotation and a dilation, respectively, which arise in the evaluation of the trace formula corresponding to the conjugation procedure in order to obtain a convenient fundamental domain F\7/(Dl). I) is the volume of this fundamental domain.
6.3 Discontinuous Potentials
217
6.2.4.11 The Selberg SuperTrace Formula. [51,428,447] ,4(2")/r162g(u)g(u)
[h(89 + i p (B))  h ( 8 9 +ip(F))] _ ,,=o
du
{7} k=l
2sinh
.
2
(6.2,197) The test function h is required to have the following properties: h(1/2 +ip) E C~ h(p) vanishes faster than 1/[p[ for p~ +0% h(1/2 + i p) is holomorphic in the strip ]~(P)I 0, and 9(u) is the Fourier transform of h(1/2 + ip). (B) and (F) denote the bosonic and fermionic eigenvalues of the Laplacian [] = 2(y + O0/2)DD, D = OOz+ c3o on the supersymmetric extension of the Poincard upper halfplane. 0, 0 are Grassmann variables. See 6.2.4.8 for further notation, and compare 6.16.3.2. 6.3 Discontinuous Potentials 6.3.1 Motion in HalfSpace.
6.3.1.1 Dirichlet Boundary Conditions. [144,196,316,317,340,439,448,613,722, 785,801,828,865] ~(t")=="
~'(x>o) t~
~
, x2dt
=(t,)=r,

m xV/xT~ ih~exp
= zv/~Tiz "
171 / t2
 2]~(x + z "2) II/2L]~ ) kdk J1/~(kx")Jl/2(kx') e i liTk2/2m
(6.3.1) (6.3.2)
The Green function is given by x(t")=z"
~ fo ~~dT 1
eiET/h
m
= ~ ~E[exp(
f :DID)>o)X(t)exP\2l i(imft''x2jt, dr) x(t')=x' ~[x"x'[)exp( h
x/2mE,x,,+x,,)] h
" (6.3.3)
218
Table of Path Integrals
6.3.1.2 Neumann Boundary Conditions. [448]
x(t")=="
i o~..,o>~,,>ox~ (~, s: ~'d,)
=(t')=='
m =vGTZ
m~(x'2 + ~"~) Sll~
i hTexp
=~
L~
2 i hT
\ihT]
(6.3.4)
kdkJ_ll2(kx")J_ll2(kx')e ihTk212m
(6.3.5)
The Green function is given by i
oo eiET/n L aT
=(t")==" i
DIN>)~
( im t" ) ~/t'
xadt
=(t,)==,
=,~[~ 1
m
,b)(• )
+ O(x>  b)O(b
1 ~ x x_~b)O(b r)MEI2r~'x/'Lhr< "(6.4.3)
The timedependent case has the form [126,251,320,402,578] r ( t t s ) = r ''
r(t')=r' _
m rv/~ ihr/(T) exp
[im (((T) r"
[~ \~~
+
0(T)r,,2~l Ix (mr'r"~ ~?(T) ]J ~,ilirl(t),]
(6.4.4)
The quantities r/(T) and ((T), respectively, are determined by the differential equations /~+ w2(t)r/= 0 , r/(t')  0 , //(t') = 1 , (6.4.5) ~ +,~(t)r
= o ,
r
: 1 ,
~(t') = o .
226
Table of Path Integrals
6.4.1.1 The Repelling Radial Harmonic Oscillator. [664] (A > O) ,(t")=r"
f r(t')=r' 
Dr(t)#x[r2] exp
[L 
ihsinhwTexprnrv/P7 W
Jim [e' (§ ~,~)dt] [2h Jr, mo)rt r/t
~' ~t r " ,2
+r.2)cothwT]lx(i~T]
(6.4.6)
+ A + iE/hw)]  ~ 1 /~t dE IF[892rwliF2(1 + A) 12 eiET/hTTrE/2wA
f m~ ,,~.. f i.~ ,~ )lVl_iE/2fiw,A/2~~r ) .
X M+iE/2hw,X/2~l~r
(6.4.7)
The Green function is given by r(t")=r"
ifo~176 ~ dT eil~T/n / l)r(t)ltx[r2lexP [ i2hm [Jt,t'' (§ +w2r2)dt] r(t')=r' /im~ 2"~ [iraw 2"~ F[89 + A tiE/hW)]W_iE/2tgo,A/2,_.._~r~,M_iE/2ruz,M2~>) =
i hw~/'(1
+ A)
w
(6.4.8) 6.4.1.2
(~ > o)
The 1/r 2 Potential. [26,96,191,274,281,467,544,609,640,642,771,785] r(t")=r"
f
Dr(t'#x[r2]exp(~~hh/: ''i'2dt)
r(t')=r'
]Ix (mr'r"~ \~]
= rx/~Tr"i~m exp rimcr,~ [2~" + r
=~
kdkJ~,(kr")A(kr')e ihrk~/2m
,
(6.4.9) (6.4.10)
The Green function is given by r(tJl)=rtt dT ~ifo~176
/ r(t')=r'
I.
t"
:Dr(t )#x [r2]exp /lrn ~t /'2dt)
2m. ~/=727;
(6.4.11)
6.4 The Radial Harmonic Oscillator
227
6.4.1.3 Attractive 1/r 2 Potential. [447,696] (n > 0) r(t")=r"
f
r(t')=r'
+ 288~j at] Dr(t'exP[hJt:"(2 § h~~2mr =
r~/'~'~r"~o~ ~  ~ K i ~ ( _ i k r " ) K i ~ ( i k r ' ) e  i h k ~ T / 2 m
(6.4.12)
6.4.1.4 The DDimensional Radial Free Particle. [33,464] r(t")=r"
r(t')=r' =~
m
iff~
Jim, ,2
L2~tr + r"b
]
(mr'r"'~
It+ 0;2 \  ~  ~  ]
(6.4.13)
The Green function is given by
ifo dTeiET/" f
Dr(t)Pt+ 072 [r2] exp ~im
§
r(t')=r' (6.4.14) 6.4.1.5 The DDimensional Radial Harmonic Oscillator. [402,464,771] r(t")r"
r(t')=r' = ~
rnzo i h s i n w T exp
rnW (r'2 + r"2)cotwT I t D2 ~ +w ~ i h s i n w T J
.
 2ih"
(6.4.15) The Green function is given by
ifodreiET/h
/ Dr(t)/h+ o;2 [r ~] exp ~imr(t')=r'
~JJ w E , , , . o  ~ t ~ ) i tT~@ T 0 ~ff/2) 2'h'~J~t*'r 2
)
r>
(§
_ w2r2)dt
M2__f_=,89 o;2 )
r< (6.4.16)
228
Table of Path Integrals
6.4.1.6 The Potential V(r) = (h2Vo2/2m)(a/r  r/a) 2. [421,771] (Vo > O) r(t")r"
m.2 / ) r ( t ) e x p "( ~i/ , t" [~r
/ r(tt).~r
2]dt~
h2V~ a  r ) 2( r n
] J
I
= ia sin \(nV~ ma
I ~
]
xexp [iliV~
iasin (nV~ \ ma
 Vo.,2 2T~a(" +,"2)r
]
(liVOT~
\~j
/] '
(6.4.17)
6.4.2 M o r s e P o t e n t i a l . [26,129,186,208,271,382,423,528,741,775,871]
(Vo > 0)
{~
~('")=~"
~
dT
Dz(t) exp
t,, jt, [ 2
 2m ( e2~ 2ae~
dt
z(t,)=z, = mF(~I + v r ~ / h  aVo) e_(#+x,,)l 2 h ~Vor (1 + 22v~~~) • W~vo,C:z~/h(2Vo e~>)M~vo,CZV~/h(2Vo e~ O)
i Dg(t)g i D(o 0) x(t")=x" f Vx(t) x(t')=x' xexp
~
x2_~
 2rrihsinwTmw ( X
3" n6~lo
~w ( z i  z j ) 2 +
m (zi  xj) 2
at
3m )1/2 exp [ 2ihT" 3m (R,_R,)2](4sin3~,sin3
32x+I(n + A+ 89 F2(A _[_1~)"'(x+ C,r~ 89 F(2A + n + 1)
+n+89 wTll3(;
3~')C(X+89 (cos39'' )
~'K~ ~t rt t
• exp Lri.  ~  ( r ,2 +
cot r,,2)
.
(6.4.29)
6.4.6.3 Three Particle CalogeroMarchioro Model. [579] (x = (zl, z~, x3) as before, xi+3  zi, i = 1, 2, 3, ~o lies in the sectors nrr/6 < ~o < (n + 1)rr/6, n = 0, 1,..., 11; A,# > 0) x(t")=x" i=1
x(t,)=x, _i_
/ 3m = 3 ~ )
h 1/2
exp[
4
m (zi  zi+l) 2
'}3
(zi + Zi+l  2zi+2) 2 J J dt )
3m . ,,
row, ,, + r,,2) cotw T] I3(x+,+2, +1) ~,/ims~~n " ) x ihsinwTrr~exp [~kr[i hw r ' rwT (6.4.30)
6.4.6.4 Three Particle Singular CalogeroMarchioro Model. (x = (zl, z2, x3) as before, zi+3  xi, i = 1,2,3, ~o lies in the sectors nTr/6 < ~o < (n + 1)rr/6, n  0 , 1 , ,11;A,a>0;N=N+A+ 89 A2 = 9 N 2 m2~2/h4N2, aN = 1/3aN) x(t")=x" x(t,)=x,
Table of Path Integrals
232
+ . o.,+.,+1_2..+2)])d, . . . .
+
(zi
m
 xit1) 2
:~C~~) ~m (R [ 3m ~112 oxp[ 2ihT
rir..~ ,2+r,,2 )
3
~gi  Xi'l1
Z ~b(~)( r 1 6 2 _ n') ~] Nero
,,
x ihsinwT exp [~(r
~}~
v~r 2
cotw
m] (rn,or'r"'~ IA \ i ~ T )
'
[9~+~2r(~+a+l/2)r(i~N+~+l/2)]
 F(2A + 1) [
N'~
(6.4.31)
~/2
F(/~    ~ =  A  Y
x (2sin 3~) x+1/2 exp [3i~o(i ~N  N)] 1 x2FI(N,A+~+icrN;2A+I;1e 6i~) .
(6.4.32)
6.3.6.5 Modified Three Particle CalogeroMarchioro Model. [579] (x = (zl,z2, x3), xi+3 = zi, i = 1,2,3, ~ lies in the sectors nTr/6 < < (n + 1)1r/6 , n = 0, 1,..., 11, A, B > 0, ~ = +3x/A + B + 1/4,,X = +3x/A  B + 1/4, A = (~ + A + 21 + 1)/4) x(t")=x"
i 7)x(t) exp x(t')=x'

XiF1) 2
i=1 .]!
,,2(
= 3 \ ~ ] 3m ~1/2 e x p [ 
A
_ _
m
~
~2(Xi
 E
B ~i + zi+x
__

2zi+2~1 . I
(xiXi+l) 2
3m ,, 2ihr/R  R')~],~,/,")~,/,'/
ri,~, ,~ .,,~)
X ihsinwTeXp[~tr
+
] (.~r'r"~
cotwT IA i h s i n w T ]
6.4.33)
~1(~o) = [ (x + A + 2n + l)n' F(x + A + n + l) ] 1/~ r(~ + n + 1)r(;~ + n + 1) t7)/sin"x~+x/2/ 
. \x+1/2
(6.4.34)
6.4.7 T i m e  D e p e n d e n t Centrifugal Potential. [170] (m(t)g(t) = K = timeindependent) r(t")=r"
r(t')=r'
6.4 T h e Radial H a r m o n i c Oscillator
~/mtm'iJ'p'rtr"
233
( ~lm#m'/J#[~" r%'~
{'[ (m'b'r'~+m"b"r"2)
x exp ~h
cot(/'/)+
rn #r #2 _ r n  r  2
+
m"'~"""~]}77 .1 ' (6.4.35)
where s(t) and p(t) satisfy (q(t) = ~
L
i~(t) ~s t ,, ,.~i"~ ()
1
)
~tt(t] _ ~(~) + (w2(t)  i~(t) ~(t))
~(t) + \ ~ ( t )

and ~ = 89
+ 8mg/h~  V2K/h~ + 1/4.
=3(0'
1
~=(t) ' (6.4.36) with boundary conditions s(t') = 1, D(t') = 0, s2(t)D(t) = 1, and D(t') = 1, "
6.4.8 A h a r o n o v  B o h m Effect. [7]
6.4.8.1 AharonovBohm Effect in the Plane. [79,86,87,289,290,362,387,515, 529,574,575,613,629,650,651,709,812,827,877,880,886,925,926,927]
Z d~ r(t')=r'
~(t')=~'
xexp
g
(§
m
rim
27rihT E ei(~'+l)(~"~')exp [2~(r vE~"
dt
,~
]
{'mr'r"~
+ r ''2) I H \ ~ ]
(6.4.37)
(f = e~/2rrcli, 4~ = magnetic flux). The Green function is given by [877] 1
sin ~rf m ~" r h 2
x'
J~
ds K0 ( v = ~  ~  ~ )
efE('i+''r 1 ~ e s+i(~p''~'') ' (6.4.38)
where the value in {.} is taken depending on whether (~"  9') e (zr, rr), (Tr, 27r), (2~r, ~r), respectively, and R~(s) = r'2+ r"2+ r'r" cosh s. Note that the claim of Ref. [83] to evaluate ringshaped topological defects via toroidal coordinates cannot be seen as correct.
234
Table of Path Integrals
6.3.8.2 AharonovBohm Effect with a Radial Harmonic Oscillator and a Magnetic Field. [87,166,168,172,387,529,771,886,925] (w = eB/2mc, ~22 = ~2 + ~ , y = e~/2~hc, ~ = x/("  f)2 + g, g > o)
,(,,,)=~,, ~d~
f
~(,,,)=~,, f
7)r(t)r
~(t,)=~,
(
~"
)
1)~(t)5 ~  ft: (adt +wT
~,(t')=v'
xexP{hft:"[2(§162
dt}
= 2~.ihsinDTe'f(~"~'+~~ x E ei~'(~~176 ~,e~
exp 
+
cotDT
( mDr'r" "~
t,~ \ihsin 12T] "
(6.4.39)
6.4.9 A n h a r m o n i c P o t e n t i a l s .
6.4.9.1 Anharmonic Radial (Confinement) Potentials. [659,865] (L = (2l + 1)/(p+ 2), p > 2) ih fO~ dt
f
W(t).,+l/2[r2lexp
~im
(§ _grP)d t
r(t')=r' 4m rv/~ft. f 2 2 ~ g (,+2)/2~ f 2 ~mgr(v+2)/2"~  h2  P 7 2 1 ~ L ~  P ~ V Y r > )~L~P"~V'~ < J(6.4.40)
6.4.9.2 Modified Coulomb Potential (Conditionally Solvable Natanzon Potential). [449,450,452] i fo~176 eiET/h r(t")=r" r(t')=r ~ _
1,2
\ 4h(r')h(r")] ~ r (  ~ )
[~ftf"(
)]
6.4 The Radial Harmonic Oscillator
235
)
4~1=1E']
tvh V(r)=
h2 g2h2 + glh3
h2 [ {h"'~ 2
h""~
(6.4.41) (6.4.42)
9
Here R = a2h 2 + o'1 h4, and the function h = h(r) is implicitly defined by the differential equation h'(r) = 2 h ( r ) / v ~ . Furthermore we have abbreviated (w ~ = cq E/2m) 1E ~ v = ~ + wh '
E~
1( = 4 ~
h2g2 h4g~ )  ~m + 16m2~1E
(6.4.43)
Table 6.6. Confluent conditionally solvable Natanzon potentials n(h)
V(h)
Range
R=alh4{a2h 2
h2 glh3 + g2h2 + AV(h) 2m R(h)
h>0
R=4h2,h=r
~m(glrtg2)
h2
R = 16h 4, h = V~ R  O'lh3 at or2h2
+ r
r>0
r2
h 2 glh 4 + g2h2 q AV(h) 2m R(h)
R = 4h2,h = r R = 9h 3,h = r 213
r>0
h>O
glr 2 +g~)
r>0
h 2 ( r213 5) ~gl + 92r ~/3  ~r 2
r> 0
6.4.9.3 Radial Confinement Potential (Natanzon Potential). [449,450,452] i L
oo e i ET/Ii r(t")=r" [ i t" dT i Dr(t)exp [ ~ t , ( 2~2r(t')mr'
=
4h(r')h(r")
m
r(u)
1 V(r))dt
J
236
Table of Path Integrals

(,,..,x
(,,...,
E)/Dv(~hm;:E)} R(,)
2,.
+~
3 g
27
'
(6.4.44) (6.4.45)
"
Here R : o'2h2 + uih a, and the function h = h(r) is implicitly defined by the differential equation h'(r) = 2h(r)/v/R, and (w = h v ~ / 2 m )
1 E' u~+~ ,
E'
1( mo'~E 2 = ~ u2E+ 2h2g~
h2g2 ~mm]
(6.4.46)
6.4.9.4 Sextic Potential  QuasiExactly Solvable Model. [458,659,903] (A 2 = (/3~ 29)/16) 
~(t")=~"
~/o~a~~(t,)=~, i ..I,~ X
{,
exp ~at,
= (XtXt')I/2 4
(k'

+ s) w'x
"~m \
~
/'v(1 ..~ )~ + ~k2/Svtl.~ ) r(1 + 2.x) 7714O 4
~'Ya9 4
(6.4.47)
6.4.9.5 PowerConfinement Potential. [71] (x E IR2) ih
dT
/
Dx(t) exp g
x2_
_
x(t,)=x,
= v~E g
2~
dv~
(r'r"W~e(1 + 21ul/d)
• ..~,/,,.m,l,.i/,,\ur>)..',~,/,,.m,H/,,\j < j "
(6.4.48)
6.4 The Radial Harmonic Oscillator
237
6.4.10 SuperIntegrable Potentials. [305,668]
6.4.10.1 Generalized Oscillator. (x
= (Zl, Z2), kl,2 > 0, [274,402,458,771,865])
x(t")=x"
/
Z)x(t)exp
L
+ ~'~
[2(*2 w'x2) 2m\ z~
x(t')=x, Cartesian: 
I
1 2 2
exp ihs:mnwT iH= 1 V~i~i ~
J
[~~tzi
l It
+ x_,,~. i )cotwT 1 Iki\ihsinwT) (6.4.49)
Polar (A = kl + k2 + 2n + 1, [139142,458]): Tt'KO
ihsinwT
E
nfilNo
~(k~'k~)(~~176
mwolO"
(6.4.50)
x e x p [  2~t0 row' '2 + 0''2) cotwT] Ix (ihsinwTJ"
6.4.10.2 Holt Potential. [274,340,402,458,771,865] (x = (z, y), x E IR, y > 0, kl EIR, k 2 > 0 , ~   z q  k l / 8 m w 2) x(t")=x" h2 k ~ _2m y2
/ x(t')=x'
/ m,..,,,,,,,,
= Vilrhsin2wT exp x ihsinwTexp
{
m,..,
ihsin2wT
88
[(,,+~.,,~)cos~_,,.,,]}
 2~(~/2+y"2)cotw
Ik, ihsinwT
"
(6.4.51)
6.4.10.3 Generalized Oscillator. (x = (Zl,Z2,x3), kl,2,3 > 0, A~ = kl + k2 + 2 n + 1) x(t")=x"
x(t')=x' Cartesian[274,402,458,771,865]: "I
/ Tl'~ iI X II \
 \ihs~nwT] j l ~ = x ~ exp [  ~ l h 'zi
(6.4.52)
238
Table of Path Integrals
Circular Polar (A1 = kl + kg. + 2n + 1, [139143,458]):
= ( i h s i n % T ) 2 ~ e x p [  r n w ' ' 2 + z " 2 ) c ~tz
\i~T]
• ~ r162162 nEl%
2~t~o
x exp
(6.4.53)
eotwT I~ 1 \ i ~ T ]
+
Spherical (A2 = 2m + A1 + 1, [139142]):
= (r'r" sin 0' sin d " )  1/~ • ~ ~(~',~).(~")) ,
(6.5.1)
~(#,.) (~,,)~(#,~)(~,) :
(6.5.2)
,
nEg%
[
,,
n'F(a4/3 + n § 1)]1/2
• (sin z)"+'/2(cos x)O+l/2p(a'#)(cos 2x) , (6.5.3) ~2
E. = ~ m ( ~ + Z + 2  +
1)2
with ml/2 = ~(fl + ~), LE =
(6.5.4)
.CRY0 +
6.5.1.e Symmetric PSschlTeller Potential. [344,526,527,742] (LE = 89 +
vff~lh, ~ > o) h
I
~(t")=x"
t" m. 2
h2 )~2 __i~
]
~(t,)=a:, m
"
t "
= ~~/sln x sin x" F(a  L ~ ) F ( L B + )~ + 1)PL:(COS x) ,
(6.5.5) ffn (x")~', (x') h (n+ +89 2 m  E '
=
~,.(z)
:
[(n+),+ 89
(6.5.6)
x/2
n!
~P;+"x
(cos x) .
(6.5.7)
6.5.2 ScarfLike Potential. [441] (IA + B[ + 1/4 > 0) ih fo ~176 ei ETllf •
/
Da(t) exp
ft,
[ 2 ~ 2  2h~~(A c~ ~  Bc~ s~n~]J dt
a(t')=a' = 4h 2 ~/sln otf sin c~"
r ( m l  L E ) F ( L E + m~ + 1)
F(ml + rn2 + 1)F(ma  rn2 + 1)
242
Table of Path Integrals ( 1  cosa' x
2
1coscJ')(m~mD/2(l+cosa'l+cosa") (''~+m2)/2 2
2
2
x 2Fl(LE+mt,LE+m,+l;m,m2+e;
1cos~
0)
r(t")=r" Dr(t)
ifo~176 ~ dT eiET/h /
r(t')=r'
L~)F(L. + ml + 1) li2 F(ml + m2 + 1)F(ml  m2 + 1) x (cosh r' cosh r") (m'm~) (tanh r' tanh r") "~'+rn~+l/2 m
F(ml

• ~(
~ + ~ , ~ +~, + 1;~,~+
1; cos~~._O) 1 9/m sinh(~r[ml + m2D/2R ~(+)(x) = F ( I + ml 4 rn2) hlsinr(ml+LB)[
• 2F1 (ml 4 LB + 1, ml  LB; 1 + ml + m~; 1 4 tanh z) .
(6.6.18)
6.6.5 T h e W o o d  S a x o n
P o t e n t i a l . [424,654] (ml,2 as before, V0 > 0)
z(t")=z"
~
dT eiET/li x(v)=~' 2m r(rnl)r(m, + 1) h2 F(ml + m2 + 1)F(ml  m2 + 1) 1 •
•
tanh
2
1
tanh ~~
2
1
tanh
2
< ml,ml + l,ml +m2+ l; 1 + t2nh ~ )
tanh l+tanh~
2
m1+~2 2
248
Table of Path Integrals
X2Fl(mi,ml+l,mlm2+l; f0 ~ !/t(+)%"~'r'(+)*' ,,
h2k2/2m  Vo  E
~(i)(~) =
1tanh~~) 2
(6.6.19) (6.6.20)
'
1 x/m sinh(r]ml • m21)/2 r(1 + ml • m2) h I sin Zrml [ 1 + tanh ~: (r~+m~)/2 1  tanh x 2 2
x2Fl(mt+l,ml;X+ml•177
6.6.6
(6.6.21)
9
The M a n n i n g  R o s e n Potential. [116,424,613,670] oo
r(t")=/'
i~ooti dTeil~T/a
/
ri t" /:)r(t)exp [ h ~ t, ( 2 ~ 9 + A c ~
)
]
sinh 2B r dt
r(t')=r' m F ( m l  L E ) F ( L ~ +ml + 1) ti2 F(ml + m~ + 1)F(ml  m2 + 1)
2 x
cothr'+l
2
) (m1+m2+l)/2
cothr"+l
(cothr'  1 c o t h r "  { ) (m~m2)/2 • \ coth r' ~ 1 coth r"
x2F1 x2F1
(
cothr>  1)
LE+ml,Lm+ml+l;mlm2+l;cothr>+l
Lm+ml,Lm+ml+l;ml+m2+l;cothr 0, Lm =  89 + ~/2m(A  E)/2, and m1,2 = ~
1 + W
•
, /  2 m ( A + E)
(6.6.24)
The wave functions and the energy spectrum of the bound states read [0, 1,...,n O) ~(t")=~" 1
~ff
f
o(t")=o" /)r(t)sinhS r
~(t')=~'
x exp h2 2mR2
{
~(t")=~"
f Z)d(t)sinO f /)!o(t) ,~(t,)=o, ~(t,)=~,'
Rs[4"s +sinh s v(Os +sin s O~b2) ws tanhS v]
[(21
1 1 1 + sinh2~ ~
11) ]})1
k 2  ~ k32 \~'m~ ~ + cos s ~o + k~ ~ coss 0
1 4
dt
6.6 The Modified PSschlTeller Potential
255
= ~ga(sinh 2 r' sinh ~ r" sin 0' sin 0") 1/2 n E]~To
, " ''r,'(='xN''~)/~''~ , .~' , Ne  i T E N / h )2 ~""')(vg")~"k')(v~') Z 'r'('xa''~)/~' N0
mE]'qo
Jr
/7
}
dk~O(X"A)(TH)e(X2'X)*(T')eiEk/t~
.
(6.6.57)
The energy spectra of the bound and continuum states have the form
E N 
[(2N+I_Aa+A=)2_I]+.~_w m
h=
2mR2
2r~2 ~
(6.6.58)
h2 ~rr~2R Ek  2mR= (k= + 88 + 
2
,
_
k>0.
(6.6.59)
6.6.14 M o d i f i e d R o s e n  M o r s e P o t e n t i a l . I ( C o n d i t i o n a l l y S o l v a b l e N a t a n z o n P o t e n t i a l ) . [449,450,452]
r(t')r'
s
= h~ \ z(r')z(r") )
2 ~
•
v~') + 1
• (1+,/~ 1  x/z(r')
+ m2 + 1 ) r ( m l  m~ + 1)
) (.,1+,~+1)/2
~ + 1
1_+ ~ ' / ( . . ,  . . ~ ) / 2 1
]
/
x 2F1 (  L E + ml,LE + ml + l;ml  m2 + l; k
17 r
• 2F1 (LE q ml, L~ + ml + 1; ml + m~ + 1; 1 + ~ \
'
(6.6.60) h2 ~z(1 z) + h0(1 z) + hlz 1/2 ~ ( (z,,~ ~ e,,~ + gg~ 3 \ z, /  2  7 ) V(r) = 2.~ n(z) (6.6.61)
256
Table of Path Integrals
Here R(z) = boz + co, and z = z(r) is implicitly defined by the differential equation z' = 2z(1z)/v/ ~ . The variable z varies in the interval z 9 (0, 1), and LB = 89  hi  2m(bo + co)E/li 2  1), ~/2 = ho + 1  2mcoE/]i 2
m1,2 = V/ho + 1  2mcoE/li 2 4 ~1 V/h1 + 1  2m(bo + co)E/h 2
6.6.62)
6.6.15 M o d i f i e d R o s e n  M o r s e P o t e n t i a l . II ( C o n d i t i o n a l l y Solvable N a t a n z o n P o t e n t i a l ) . [449,450,452] r(t")=r"
i [oo e i ET/tf J0 d T r(t')mr'
2m f R(r')n(r") ~ 1/4 F(rn,  L~)F(Lv + ml + 1) = li2 \ ~ ] F(ml+m2+l)F(mlm2+l) x[l(
•
1 + V/I l z ( r ~ ) ) "
( 1+ V/llz(r,,))]
(m~m~)/2
l+V/lz(~,)1+~~(~,,)
• ~F~  L ~ + m ~ , L . + , ~ + l ; m ~  m ~ + 1 ; 2 • ~F~  L ~ + m~, L. + mi + 1; ~1 + , ~
1+ V/1 z< (,')
+ 1; (1 + , f i ; 7 >
(~))~
' (6.6.63)
h2 3z(1z)+h~
V(r) = 2m
+~Ii2 \( "( z~; " 1 2 _2_7)z"'~ . (6.6.64)
Here R(z) = boz + co, and z = z(r) is implicitly defined by the differential equation z' = 2z(1z)/y/R~ . The variable z varies in the interval z 9 (0, 1). Furthermore Lo = (x/h0  hi + 1  2mcoEh ~  1)/2 and 1 h o + hi + 1  2mcoE/h 2 4 V / i  2m(b0 + co)E/h 2) ml,2 = ~V/
(6.6.65)
6.6 The Modified P6schlTeller Potential
257
T a b l e 6.7. Hypergeometric conditionally solvable Natanzon potentials
R(z),z 9
V(z),
R = bo z + co
82 3z(1  z ) / 4 + h o ( 1  z ) + h l z 112 + my(z) 2m R(z)
R = I , xEIR z
89 + t a n h x )
82( 2m h o + l
8~ (
R=lz,r>O
3
3 ) 4(1+e_2~)2
h
coshr'~ 1s i  '  ~ r J
ho + 1 + 4 sinh'~r+
z = 1/cosh 2 r
82 3z(1  z)/4 + ho(1  z) + h,(1
R = b o z + co
2m
R=I, x61R + tanh
ho  3/4 hi l+e_2~ + /1+e_2~
82 ( h o + 3]4 2m ~, ~  ~ r + hi coth r + 1
z = tanh 2 r
89
z'J
)
R=z,r>O
=
A V = 2'mi \ z ' }

Z) 1/2
R(z)
82 (
ho3/4
hie ~
3
~) 2m ho + ~ 1 + e~~+ ~
R=z,r>O
82 ( h o + 3 / 4
.
+ AV(z)
coshrXX
z = tanh 2 r
R=lz,r>O z = 1/cosh 2 r R = ao2
+ boz
R=zr>O z = tanh 2 r
82 ( 3 +hlcothr) 2m h0 + 1 + 4 sinh2"~
82 f z ( z  1 )  3 ( 1  z ) / 4 + h l Z 2m
3/2
R(z)
+ avo)
82( f+3/4 ) 2.~ ht tanhr cosh~7~1
R=zZ, r>O
8~ (
z = 1  e 2r
2m
y+3/4 f + 1
h,
3 )
27 1 e + x/1  e 2r
4(1 _ ~  2 r ) 2
R = 4z(1z) h~ sin(~/2) cos2(~,/2) 
e (o,~) R = aoz 2 + boz
h2 (h
z = tanh2 r
2m \ ' ~ ;
sinh r
R = z 2 , r >O n = 4z(1  z) z = 89 cos~)
e (o,~)
+ 4 cos2(~/2)
8 2 fz(z  1)  3(1  z)14 + hxz3/2~/1  z 2m
R=z,r>0
z = 1  e 2r
y.+
2m
f+1
h2F
~m h, t a n ~ 
n(~) f + 3/4
+ ~v(~)
'~
cosh~ + 1}
i+3,4 h,er
1   e 2~ +
f+
)
4(1 +~2~)2~
x/1  e 2~
3) 4(i _~2~)2
0 3+ 14cos2(~/2)
258
Table of Path Integrals
6.7 M o t i o n o n G r o u p Spaces a n d H o m o g e n e o u s Spaces 6.7.1 G e n e r a l F o r m u l m .
6.7.1.1 Motion on a Group Manifold. [104,262,263,447,454,618,679,726,776] gtI t l l ~)=gII
 ~ " ~t' (gl(g _1_i ~ g ) , g  l ( g + i ~ g ) ) d t g(tQ=g'
/,
/ dEt dtx t (9) e i EtT/h _ (2r)p+t dv/d~D 
(6.7.1)
e i hnT/48
1I~eR+ (27rihT)"/2 •
~
(a, ~ + 2Try) 2 sin( 0, r E IR)
e(t')=e' r(t')=r' /rt ~dv ei ,,(~"r') J~o~ kdk
eink2T/2m (6.7.~2)
6.7.1.10 Elliptic Coordinates in In, (1'1). ,~(t")=,,"
f
b(t")=b"
7)a(t)
,,(t,)=a,
f
:Db(t)d2(sinh2asinh2b)
b(t,)=b,
[im d2 [t Jt'
• exp [ 2h
1Jo
= ~ 81r
•
[447,470] (a E IR, b > 0, d > 0)
kdk
"
(sinh2 a  s i n h 2 b)(h2 _ b2) dt
]
d u e ~vihk~T/2m
Mei.(b, jt4'(3)2 ,'",v *(a'; ~ ) e ~ ' ' rJ t * ' ~ ) M i ( ~ ) ( a''; kA~
I
(6.7.13)
262
Table of Path Integrals
6.7.1.11 Spheroidal Coordinates in ]R(2'1). [447] (~,r} > 0,~ E [0, 2~r),d > 0)
r
o(t")=o" ~D((t) / ~Dr/(t)(sinh2(sinh2rl)d3sinh(sinhr/
/ ((t')=('
rl(t')=r/'
x exp
v(t")=v" f 79~p(t) ~o(t')=~o'
d2(sinh 2 ~  sinh 2 r])(~2 _ ~2) _ sinh 2 ~ sinh 2 y r
8md2sin~:(sinh2rl] dt} ei ~,(~o" ~o') f o o
d.
1o .sio
J0 2Ve
v "eosh r/. .,r. 2 aa. n)rsi._lp.(cos ~.. . h r/,"k2d 2) x Ps i.1/2( ~,(3) .. , ~,( 3 ) 9 , • S i._l/2(cosh~ kd) S I .  1/ 2(cosh ~ ; kd) .
(6.7.14)
6. 7.1.12 Summation Formula for Path Integral on a Quotient Manifold (Mirror Principle, Method of Images). [340,541,676,679,828] T) =
K,.(,",
T) .
(6.7.15)
7EF
6.7.2 Motion on the DDimensional UnitSphere. [25,26,104,105,136, 257,282,357,358,444,447,464,468,528,612,613,653,660,678,679,762,763,787,826] n(t")=fl"
f "DrY(t)exp { h ft:" [ 2 ~ 2 + h2 (D  12(2  3) ] dt } n(t,)=n, = ~
~_, S~"(ff)S~'(a")exp
I E N o t~=1
_
1 ~2I+__D~/2(n) l~~o ~ D

(6.7.16)
+ D  2) 217I
C[r_(cosr
'
l(l+D2)
(6.7.17) S~(a)
1 Ns(._I )
D3 k=0
(6.7.18) E S(O1) = h2 l(l + D  2) , l 2m
(6.7.19)
6.7 Motion on Group Spaces and Homogeneous Spaces
263
where ~2(D) = 21rDI2]F(DI2), I = mo > ml > ... > roD2 > O, N S(o') D2 2~rllk=l E k ( m k  l , m ~ ) and Ek(l, m) =
+ m
7r2k2m(D2)P(l Dlk

k+ D

1)
(6.7.20)
(l + .2~)(1  m)!P2(m + DZk~ 2 s
and cos r
denotes the quantity defined by (2.7.3) For D even one has
fi(t")=n" fl(t,)=fi, 
ri hT(D2)2J] (~ cosr 8m
1
(27r)D/2 exp I.
D2
T~ (~
 2~m) (6.7.21)
The Green function for D even has the form II(t")=tl" i L ~ 1 7 6 eiE'T/h
:Dfl(t) t'l(t')=.fl'
2 E + h ~ ( D  2)2i4m
= g
x
2~ dcosr
,)
sin (Irvl2mEIh~ + ( D  2)~14)
' (6.7.22)
m r(a+Of~l)( 1 ) D~'3 3o  2h2sin~r(a+ 89 r ( a + 5~.._~D) 2~rsin ?(,,,,) P,'F[cos(r
~r)] (6.7.23)
(a 89 + v/2mE/h 2 + ( D  2)2/4); and for D odd ~(t")=n"
iLCCdT h eiEWli
i ~Pfl(t) n(t,)=n,
Table of Path Integrals
264
m F(a+~)( = 2h2sin(arr) r ( a + L~_)
D3
1 2~r sin r
) 2
3D
PAT(  cos r
6.7.3 B i s p h e r i c a l C o o r d i n a t e s . (p = l + ( N  2)/2,
v
=
. (6.7.24)
A + (M  2)/2)
[1041 r(t")=r"
I
0(t")=#" :Dr(t)r N+Mi / :Dtg(t) sin N1 t9 cos Mi
o(t,)=o,
r(t')=r'
i[~N_l(ttl~i__flN_ltt , ,o
n.
fl~I (t')=fl~l'
flaM' (t ')t'l~Ml'
.ox. { i
MI
t"
II

(t)_n~
MI
II
sin2 tg~2a + 0:
x(t")=x"
iL~d T eiEr/h f
= ~
v(t")=v" 7~x(t)
x(t,)=,e
f Vv(t) yO1
v(t')=v'
, \ ~ ~m(D1)(D3) [~i....~.+, .
•
~(2~') r 1 6D2 2
)] dt
Ir~ f0 ~ ~ ~kEf~ D~ ~. e,, 0,12 E s(D2):
~(t")=~" n(t")=n" = ~ L~176 eiET/h 7 l)r(t)sinhD2 T i D"(t) r(t')=r' • exp
n(t')=12'
(#2 + sinh2 r O 2) _ h 2
'~ ] dt
(6.7.33) Equidistant, r l , . . . , 7"o1 E ll:t~ ~,(t")=~ I'
ei = + l : rD1(t )=r~_i II
= ~
~i0~ d T e iE:r/h 9,(t')=~ i I D n (t) cosh ~
{is,,[ {is,,[
Tl
~
.
n...
II
f ~Dl(t')=~5_,
.
.
.
:o~o 1(t)
267
6.7 Motion on Group Spaces and Homogeneous Spaces
h2(
1
i
~ ~ + ...+ 8m (D  2) 2 + c~ 2 n cosh2 v: ...cosh 2 tO2
)]} dt
= (coshD_2r~ coshD_ 2 r 1,, • ... • cosh r/)_~' cosh r~_ 2):/2 D2 H E
[ dk~ x Jr, 2~ eik~
j = l ej=4
Pii:~_x_l/2(~Dl_j
1 [oo dkj kj sinh 7rkj Jo (cosh 2 lrkj_: +sinh27rkj)
:
Tf)_l_j)Pi'klk_~_l/2(eD_t_j EkD2  E
tanh
X
General expression for the Green function: m ( e  i r r ~(D3)/2 rch2 k.2~dl~d/ 1/2i~/2m(EEo)/a (c~ _
_
tanh r~_l_j)
(6.7.34)
'4))
"
(6.7.35)
!
(x = {xi} = (X:,...,ZD_2), r 2 = E~'_~2z'~). In the equidistant system we identify EkD_2 = Eh with Ek = Eo + h2k2/2m, Eo = h2(D  2)2/8m, the wave functions H(,~)~,(f~) (6.7.29) from Sect. 6.7.5, and, e.g., cosh d(q", q') =
Ix"  x'l 2 + y,2 + y,,2
2y~y.
(6.7.36)
6.7.6 PseudoBispherieal C o o r d i n a t e s . ( N = 0, 1 , . . . , N M < 8 9  # 1), # = l + (N  2)/2, v = A + (M  2)/2) [104]
r(t')=r /)V(t) sinh NI "rcosh MI
T
r(t')=r' a~"~Nl(t'll~[~NItl\
~'~N1
,,
J~
(t l'~jn,.N
• exp { i / : "
~~M1(t II )=~~M1 H
lt
n aM   1 ( t )t  ~ a
M ll
[ 2 ( / " + sinh2 r " ~ + cosh2 r 0 ~ )
8m (N + M  2) 2 + c~ 2 r N
= l,XE]tqo N
ct N (aolsi
t
M M
sinh r}J
J
268
Table of Path Integrals NM
'~"~ t~.,h~(u,v) [~.,h,r,(u,v) *eiENT/h X (~N~==O ,.].N '  /~:N
+
dkO(mv)l~'q~r~(mv)*t.r~eiEkT/h (6.7.37)
2N!vF(v N)
1/2
x(sinhr)l+~,~p(~,uu2N1)(isinh)r.)
(6.7.38)
cosh 2 T
EN = 2''m (u  p  2N  1) 2 
4
(6.7.39)
'
O~"'")(r) = N(m~)(cosh r)l+"(sinh r) 1+~~
•
l+v+pik 2
'
l+u+p+ik 2
h2 ( (N+M2) E k = ~ m k2+ 4
; l + p ;  s i n h 2r
)
,
2)
(6.7.40) (6.7.41)
"
Here kl = 89 + v), k2 = 89 + p), and N (m~) as in Sect. 3.4.5.2. 6 . 7 . 7 The SingleSheeted Hyperboloid. [8,442,447] (n  0, 1,..., NM
0,~ = ele2x/ml2E/h,a
6.8.2 K r a t z e r
= h2/mele2)
[864]
i
dT eiET/h
h
:Dr(t)exp ~
,
~
+ ~

r
2m
~ dt r2
r(t')=r'
1 fmF( 89 + )~  ~)
= ~
V,~
nell/0
n+a+ 89 xexp
E, =
+
~.(,")~(r')
[
:? , f~dk ~e,(.")e;(,')
~r(n+2~+l)
a(n+~+l/2)
m(ele~)2 ~]2 ' +
2h 2 (. +
k~k(r) = F(89 X/~F(2X+
\~(n+~+
(6.8.7)
89
a(n+)~+l/2) n = 0, 1,...
" exp (Tr~ak)M i/ak,X(21kr) 9
'
(08 , (6.8.9) (6.8.10)
278
Table of Path Integrals
6.8.3 P u r e C o u l o m b P o t e n t i a l in Two Dimensions.
[280,356,514,564,613](~
6.8.3.1 Green Function.
= e:2V/Z~/2EIh,
x =
~(cos ~, sin ~) e IR2) i
~
co
dTeiET/'
Ox(t)exp S, (2x2+
/
ele2
'x'] d/
x(t')=x,  2~1 E
.e~
ei V(~"~')
~
1
II
m 1 ( 1 ) 2E I~.1~.r ~ + I . l  , ~
(6.8.11)
6.8.3.2 Polar Coordinates. [280,514] (a = h2/mele2)
x(t")=x" x(t')=x' oo
= ~ Z ~'N,(o",r162162
~
ve~Z N=I * t ak ~'k,,,(otl , ~ ,'t )e;,,(o , ~') ~ i l i k ~ T / 2 m
+
(6.8.12)
v E,Tz
[ _(_N_I.I_:~ 1/~ ~VN,.(O,~) = L~a2(N_ 89 + I.I_ I), ] • ~xp EN "
m(ele2)2" "
~'h,,.(#, ~)  V:4#o x exp
1 ~] ~ N  l v l  1 ~ ( i  ~) 1
~ ( g  89
2 h 2 ( N  89
20
N = 1,2 .... ,
'
(21vl)!
+
,~l~l
(a(N89
(6.8.13) (6.8.14)
I,,I+
 iu~o M i / a k , M (  2 i k o ) .
(0.8.15)
6.8 Coulomb Potentials
279
6.8.3.3 Parabolic Coordinates. [280,458] (a = h21me,e~) ~(t")=C
i
,7(t")=,7"
V~(t)
~(t,)=~,
f V~(t)(~ ~ + ~) o(t')=o'
x exp [~/f" (2(~r + r/2,(Al0', ~o'>,
(6.8.92) (6.8.93)
6.9 Magnetic Monopole and Anyon Systems (r"l~(E)lr')
295
gx(r", r'; E)
=
~(,,,,)=r,,
_
ih
:Dr(u)
L= / du"
2mT,/r//
r(O)=r I
xexp ~
§
h2A(A+l)
a
52mr + ~[ + ~m du

F(A + 1  p) = 2 i~72~ ~ 2) Wp,A+} (  2 i kr>)Up,~+ } (  2 i kr = A(A+I)IA >; furthermore
(.,'s~ ~(t,)=~, " 0 [215,427,465,481,624,821]: i
o0
=hfodT 
7r3
e iEw/lf x(t")=x"
f kl
f
x(tt)=x'
y(t")=y"
:Dx(t)f
f
:Dy(t)y2exp ~~ft,im
t" x2 1 Y2d~! Yf ,]
y(t')=y'
e ikx(x''x'lf~176 dk ksinh 7rk
so h2(k2 + 88
E;Kik(Iklly'')Kik(lklly') 9 (6.10.4)
306
Table of Path Integrals
General expression for the Green function [435,465,466,481]:
= ~rlim~Q_l/2_ix/~,nE/~2_z/4 (cosh d(q", q')) ,
(6.10.5)
where Q~ is a Legendrefunction of the second kind, and d(q", q~) is the hyperbolic distance invariant with respect to the group action in any of the coordinate systems on A (2). Cf. [444,447] for further details in expanding the path integral on A (2) and A (3) in coordinate systems which separate the path integral. 6.10.2 M o t i o n in t h e H y p e r b o l i c P l a n e w i t h a n O s c i l l a t o r  L i k e P o t e n t i a l . [429] y(t")=y"
x(t")=z"
i
oo
ei ETIh
~(t,)=~:,
f .....
u(t')=~'
,,, ,..,, ,, F[89
+ v + EAl~)]
(6.10.6) I/
CCdk
I
*
I
E;   f
"
'
(6.10.7) v = v / l / 4  2mE/li 2, and with ~ ( x ) and E~ the wave functions and the energy spectrum, respectively, of the Euclidean onedimensional path integral problem with potential V(x). The energy levels are given by (Ex < 0)
f.with n  0, 1, 2 , . . . , NM < /2.!(IE~l/r~
(_
8m h2
2mh2 _ _ l I _ 2 .  1
(6.10.8)
]E~J/2hw  89 and the wave functions are  2.
rryM 2

1)y
(0100>
The normalized wave functions of the continuum states with Ek = h2(k 2 + 1/4)/2m are given by
6.10 Motion in Hyperbolic Geometry
307
I h k sinh rk
  27r2y F[~ (1 + ik + ~~x)]
~k,x(z,y) = ~[ mw V
.
(6.10.10)
6.10.3 M o t i o n in t h e H y p e r b o l i c P l a n e w i t h a C o u l o m b  L i k e P o t e n t i a l . [429]
yt't"')=y "
x(t")=x"
h fo~176
:Dx(t) J x(t')=x'
:Dy(t) y2
y(t')=y'
x exp {.~ Jr:" [2 ;~2I.y2 y2
ot
}
W~r (1 + ~,),/~~/h
J
(6.10.11) =
f
dEx
,~t
[~'~ Ln=O
, y ) . , x t x , y') ]~~_~
'~"'"~*''
/. oodk ~,x,p(x", y")~,p (x', y').l + Jo
~2
J (6.10.12)
\
v = x / l / 4  2mE/h 2) with ~x(z) and Ex the wave functions and the energy spectrum, respectively, of the Euclidean onedimensional path integral problem with potential V(x). The wave functions for the continuum states with Ek = h2(k 2 + 114)/2m are
~k,X(x,y)
[ hk sinh 27rk V27r2~FL~
/ 1
a +ik + 2 h ~
)
(6.10.13) For the bound states one =
/(l~ln/~
has [n = O, 1,...,NM 2n  1)n! "2y
c
O) ~
d T e iET/~ =(t")=z" U(t")=y" x f ,x(t) / :Dy(t)  y2 exp [ i~~ t t ," ( 2 ~ : 2 + Yy22
=(t,)=='
bh y ) dt ]
u(t')=u'
m F(89  b~,  i k ) v /  ~ '
2rh
r(1  2 i k)
x L ~le du i.(~"  =') Wb~,i~(21ulY>)ib~,ik(21vlY>),
[
m F( 89189
_
 2rrh
F(1  2 i k)
2ib,~
exp [  ~
artanh
(6.10.16) + y,
• (cosh 2)2b" (sinh 2 ) 2(b' 89 a, x2F1 NM
x.
f
oo
: Jo=
(1
1
2 )
bL,ik,~b~,ik;12ik;l_cosh II
I
*
r
(6.10.17)
I
. ~,,,,,(z , Y' )~,~ ,.(z, if) ~'k,~,tz
+
dk
,U ;
du
k,~,t
Ek  E
,
if) (6.10.18)
with cosh r the invariant hyperbolic distance, k = v / 2 m E / h 2  b2  1/4. Wave functions and energy spectrum of the discrete spectrum are E,~,v=~
1 /b~,n
b2 + ~ 
]
,
(6.10.19)
 2n  1)n! ei w eVY(2vy)bvnL(2b~Zn1)(2Py) k~n,v(x,y) = V/(2by "~v'F('~i~'"~)
(6.1o. o) 1 > 0,b~ = bu/[u[). The wavefunctions with (n = O , . . . , N M = by  ~,k Ek = h2(b 2 + k 2 + 1/4)/2m of the continuous spectrum are
6.10 Motion in Hyperbolic Geometry
309
[k sinh 2rrk .. ~k,~(x,y) = V~] ~r3r3l~ I ' ( t k  b v q 89
iVx
(6.10.21)
6.10.4.2 The Pseudosphere and the Poincard Disc. [425] (bt = bl/lll, r, Ek as
in the previous example) r(t")=r"
ifoO~ dT
/
~o(t")=~o"
f
~Dr(t) sinh r
r(t')=r'
V~(t)
~o(t')=~o'
x exp { h fti" [2(#2 + sinh2 r~b2)  bh(cosh ~  1 ) e  ~
r(t")=r" i fo~ dT eiET/a = ~
/
sinh 2 r
dt
r162 :Dr(t) (1 4rr2) 2
r(t')=r'
[h;" xexpL "
1)]}
1
f
/)r
~b(t')=~b'
( § + r2'2 \2m(iZr~
r2
'
(1 r2)2 ~ dt] J
2bhl"~~r + h2 32mr ~ /
m e_2ib, {'i('* r 1 8 9 1 8 9 = 2rh \i r " r YT/
r(1  2ik)
( 2)89 (1 1 x 1tanh 2 2F1 ~ + b  i k , ~  b t  i k ; 1  2 i k ; c o s h 2
1
)
~ ,
(6.10.22) = S,
~ ,T,A~,_,, ,.,,~g,a~ 9(r', ~') ~n,l ~ ~'f" ) n,l
IE~' n=O
oo
+Z
f0
2
,,,ABr~,, ~ " ) e ~ ?
2 ,
(~',~')
dk~k'l ~ ,
(6.10.23)
Here we have = artanh
77/
'
(6.10.24)
and the coordinates ~ = x+i y are related to the pseudospherical coordinates (r, !o) via r r ~ u  tanh 7 ei(v+,~/2) (6.10.25) Furthermore denote k = x/2mE/h 2  b2  1/4. The boundstate wave functions and the energy spectrum have the form (n = 0, 1,... < NM = bt  89
310
Table of Path Integrals
, r . a . , ~ = = [.!(2b,+lt_i)r(2b,  + IZl)] '/2 :'~'ttr't~ L 47r(n+ltl)!r(2b,~) ]
• eit~(tanh2)lq(1tanh22)b'np(lq,2b'2ni'(12tanh2~)
,
(6.a0.26) = 2mr + 4 " The continuous spectrum is given by (k >_ 0, Ek as in the previous example)
, /ksinh27rk ( ~ _ k g ~ ' ( r , ~ )  ~r(lll!)v ~r .
+b,+lll ) F ( l+ik2
) bt
x ei'~ (1  tanh2 2 ) 89 (tanh 2 ) Ill 1 • 2F ~(~ik+b,+lll, 89
r
9
(6.10.28)
6.10.4.3 The Hyperbolic Strip. [425] (~, bu, r, Ek as in the previous example) X(t")=X"
i f ~ 1 7dT 6 eiET/a ~
f
Y(t")=Y"
f
DX(t)
:PY(t) cos2 y
[iLf"(~X241V2 x(t,)=x,
xexp ~
v(t,)=g,
) ] bhtanY.~ dt
cos2 ~
m eib~(y,_y,,_v) F(89+ bu 27rh •
i k ) r ( 8 9  b~
ik)
/'(1  2 i k)
( 1tanh 2
2F1
,)~ ,
+ b .  i k , ~ ,b u  i k ; 1  2 i k ; c o s h 2
(6.10.29)
= N)_~L dk vs. n r~,, ~ , y,,)ws..(x,, Y') n=O ~ En  E
+ frtdk f~ dv Cs" (X'' Y't)g~s"  E *(X''
(6.10.30)
The discrete spectrum is given by
[ n!(bu+iu)F(b~+iu.)
]x/22n_b. n)]
X eivX (ie iY
)i~"(cosY) nb~+lP(nivb~'2b~2n1)(l~e  2 i Y
) ,
(6.10.31)
6.10 Motion in Hyperbolic Geometry

=2ml

311
, n=O, 1 , . . . , N M < b ~  8 9 (6.10.32)
The continuous spectrum has the form
gZSkS(X, Y)
= N;s,ge i v X + i b Y (icos y )  i ~
(89+ i ( v  k ) , ~ +i i ( v + k ) ; l + i v  b , ; l + e 2 i
•
1
Y
) (6.10.33a)
1
~/k sinh 2rrk ~ F( 89189
N;~ = rcF(l+iub~,)V
. (6.10.33b)
6.10.4.4 The SingleSheeted Hyperboloid. [215,442,447] (7 E IR, eihT/8rnR a r(t")=r"
R2
~p(t")=qa"
/
7)r(t)sinhr
~(t')=~, x exp
[0, 27r))
~o E
f
Dg~(t)
~(t,)=~,
8mR 2 c~ 2 7 dt
R2(/2  cosh 2 r~b2)  ihbsinh r~b 
1 9 _ ~')  2~.R2(cosh r'cosh r,,)_l/2 Z e't(~" tED 1[ NM ,,/./(a) (7t(3~,b)
•
~
~
_,,b)(~,,)~(~e~,b).(e) e_ iT=. 
/~
2 L n0
f~,
+
/'it/(3)b~][TII~.r~{7I(3) 1' b~* ~
,r.k
1'
dku,~
( )w~
z
r,
(r ] e
(~(a~'b)/li }
iTE h 
.
(6.10.34)
The energy spectrum and the wave functions are given by, respectively (n = 0,1,2,...,NM
< ~(ll+
cu(~)b~ E;,
"
'
~n(~) b,
h2
b~lll

b~l 1),bt 
=
bUilD
[88
2~h2
]
+ b2  (n + 89 89 + b~l + 89
[fit + b,l  It 
b,ln
l)n!r(ll +
b,l 
b'l)2
'(6.10.35)
n)] i/2 J
x(isinhr1)
89189
x p(Itb,I,It+b,I)(isinh r) . (7/(a) b~ The continuum states with E~ ~' ' = h2(k 2 + b 2 +
(6.10.36)
1/4)/2mR 2 have the form
Table of Path Integrals
312
~n TM b~ ffk sinh 2rrk ,," '(r) = ~ry(1 + It ~1)
x r[ 89 + Pb,I + P + b,I) ik]r[ 89 + P  b , I  It + b,I) ik] (i sink2r  1 ) ik 89
x (isinh.r + 1) 89189
x2 F1 (1 (1 + It  b,I + II + b,I)  ik, (l+llb~lII+btl)ik;l+llbll;isinhr
+
.
(6.10.37) 6.10.4.5 Motion in (D  1)Dimensional Hyperbolic Space with a Magnetic Fietd. [435] (b = emB/2ch, Ay = B u = 0 gauge)
x(t")=x"
f
u(t")=u"
~x(t)f
:Dy(t) yO1
y(t')=y'
x(t')=x'
•
ff'\
Y
8~(D  1)(O  3) dt
N~
=
E E*o,(x",Y )
n,v(" " ,Yl) e
Is n = 0
dk ~ k,*,t~Xtt,Y t t)~ ,k,,,t~Xt ,Y)t~eiEkT/h
+
(6.10.38) (x = (xl,..., z D2) and similarly for b and u). The wave functions and the energy spectrum for the discrete spectrum are h2 [ (
En~'~m
v)
orn
0 'x
1) 2
+b2+
( D  2 ) 2] ~ ,
(6.10.39)
[.!(2~ 2n  1) V
x (21kly) '~" elk, v L(2~,2,,x)(21vlY) ,
(6.10.40)
with n = 0,..., NM < ot  1, ot = l~. b / I v I. For the continuous spectrum we have
h2[
Ek=~m
k2+b2 +
(02) 2]
(6.10.41)
6.10 Motion in Hyperbolic Geometry
313
e i vx /" Ok,u(x,y) = ( 2 ~ F ~ i k  a +
l '~ 2iV
/k sinh r k 2~'~
W,,,ik(2lvly)
(6.10.42)
6.10.5 Kepler P r o b l e m in a Space of Constant Negative Curvature. [60,426,460,462,463,824] (r > 0, 12 E S (D2)) r(t")=r"
/
ofi~176
vl(t")=~" Dr(t)sinhD2r va(t)
/
~(t')=~, x exp
[i~]"(_~h
n(t,)=~,
) ]
ele2, .. ~= + sinh 2 r ~ 2) + ~tco~n r  1) dt
R2
M
= (sinha'sinhcd')(~ X
,e~0 ~ u=*~5~(n")Sf(12') L e ) V ( L E + ml + 1)
m V(ml li2 F(ml + m2 + 1)F(ml  m2 + 1) (
x
1 ) }(~nl~lt~,.~,)( .]
1 l+u'
l+u"]
?Ul, ~ }(mlm2)
\l+u''
l+u"]
x2F1
LE+ml,LE+ml+l;m~m2+l;l+u
0]:
6.10 Motion in Hyperbolic Geometry • 2F1 (89 + s + i ( k 
317
k)], 8 9
+i(k
k)];1 + i / r
(6.1o. 9) Nh(.,.)
[k sinh rrk
1
 r(2k~.iV
~
[ r ( k l + k:  ~ ) r (  k , + k~ + ~)
h2 ( E~
(D2)2)
2mR5 k2 +      (   
(6.10.60)
,
x F(kl+k2+~l)F(kt+k2tr
e2
(6.10.61)
R
6.10.8 Motion in the Hyperbolic Space SU(n, 1)/S[U(1) x U(n)]. In the space $2 ~ SU(n, 1)/S[U(1) x U(n)] we have for the metric dy_~2 1 ~ _ ~ 1 (d as = + dz dz; + 7 Y k=2
+
n
z;dz
)
2
(6.10.62)
k=2
zk = x k + i y k E r (k = 2 , . . . , n), xl E IR, y > 0, with the hyperbolic distance
given by c~
d(q", q ~) = [((x"  x') 2 + y~2 + y,2)2 + 4(x]'  x~ + (x"y'  y"x')) 2] 4(if y") 4
(6.10.63) The symmetry properties of the space give rise to two important coordinate systems, in which the problem is separable, namely (n  1)fold twodimensional polar coordinates according to xk = rk cos~k Yt, = rk sin ~k
(rk>O,O 0 [435]
6.10 Motion in Hyperbolic Geometry
321
dT ei E T I h
r(t")=r" x
f
7)r(t)exp
{ i f t t " [m. 2 h2 ( ( 2 1 + m ~ + m ~  l ) ~ , ~r  ~m sinh 2r
21
r(t')=r' m r(ml  L~,)F(L~, + ml + 1) h 2 F(ml + m2 + 1)F(ml  m2 + 1)
• (cosh r' cosh r")(m'm~) (tanh r' tanh r") m'+m2+'/2
(
x2F1  L v + m l , L ~ + m l + l ; m l  m g . + l ; c o s h 2 r
) , (6.10.76) =
fOo~dk ~/:/K(Ttt)~f:/K*(Tt) ~]~~
(6.10.77)
,
(ml,2 = ~(T/+ 1 ~ / h , Lv = 8 9 21 + m2a  1), with energy spectrum
1), r/ = 2 1 + m s + m2a  1 , v =
Eg/K = h2 8m(ma + 2m2a) 2,
h2 k 2 A Eg/K, E~IK = 2m
(6.10.78)
and the wave functions are given by = #G/K(tan h
x 2 F I [ I + 8 9 ~+m2~ik)'2\l/'m'~2 + l  i k ) ; 8 9 (6.10.79a) /ksinh r k F[l + 2~
2
JV m2,~ +1 k)]F[~( 2 + 1 + i k)]
F[l + 89
+ m2a + 1)]
(6.10.79b)
322
Table of Path Integrals
6.11 E x p l i c i t T i m e  D e p e n d e n t 6.11.1 T r a n s f o r m a t i o n
Problems
Formulae.
6.11.1.1 General Transformation Formulae. [126,180,251,259,294,440,640,643, 737,847,870] (d' = e(t"), c~" = o,(tl'), etc.; we consider y = f(t)(x  c~(t))) ~(t")=#' x(t')=.' l#
1
• K(v)
~
p
;~(t"),,(t')
.
(6.11.1)
Here we have introduced the notation
y(r")=y" K(V)(y,,, yl; r", r') =
f :Dy(t)exP[hL:"(2~t2V(y't))dt y(r')=y'
] ' (6.11.2)
 F(t)z + 2w2(t)xU+h(t )
(t)) V(y,t) =  ~1V o ( x  a~(~
w2(t ) = f(t)](t)  2j2(t) f2(t)
1 ,
y(t)
e(t)
r(t) =
(6.11.3) do"
'
e~(~r)
'
(6.11.4) (6.11.5)
h(t) =
( y(t)](t)  2P(t) ~(t) + ~(t).(t)~
mi x
212(t)
(6.~1.6)
/
6.11.1.2 Explicit TimeDependent Potentials. [180,251,259,440] For the explicit timedependence we take ~(t) = x/at 2 + 2bt + c ((i = ~(tl),
r = r x(t")=x"
etc.) '
(2(t) v x(t,)=x'
dt
6.11 Explicit TimeDependent Problems
323
[ i~m ( x , , 2 ~r_ x , 2 ~ ' ~r]
= (r162
/'x" x' K~,,v~,~7, r
fti" r dt )
(6.11.7)   x,~] = (C"C')~/2exp[~i m \~I x " ~" c" c'/j f dEx~x ,'x,,, 7/7 ~P~(x,) exp ( iE~ft)" at ) 
(6.11.8)
where f dEx denotes a LebesgueStieltjes integral to include bound and scattering states ~x with energy Ex of the corresponding timeindependent problem, and with the path integral K~,,v given by
,~,v/.,,,~,;.,,/: / ./ex~
~ V~"m~'~~~/~/~d~
z(o)=z' (0.11.9)
w12 = ac b2 and s" = v(t11), where
I
1
at + b
tlt
('~
= ~7 arctan   7 
> 0) ,
tI
1 artanh at + b r
~(t") = f f " r dt
I~'1
~
t,
r,~,2 < o) ,
t In
a
t
bat+ b
(~a'2 = O)
tI
@
(6.11.10)
6.11.1.3 Moving Potentials (Extended Galilean Transformation). [276,322,440, 497] (q' = x'  It, f' = f(t'), etc.) x(t")=x"
~(t,)=~, =
exp
]"(x"
 f")
 ]'( x'  if) +
1
c ]~(t)dt )]
q(t")=q" r i t" ] • / 1)q(t)exp [~ft, ( 2 0 ~  V ( q )  m ] ( t ) q ) d t j .
q(t')=qI
(6.11.11)
324
Table of Path Integrals
6.11.2 Examples. 6.11.2.1 TimeDependent Harmonic Oscillator. [180,259,440] (~(t) = ~/at ~ + 2bt + c, /2~ = r +w,2, r,w' as in (6.11.10))
~(t")=z" x(t,)=~'
= (~'~")W~ exp [ i~m" ' ~\fx''~ (,7  x'2 (4')~ ] • exp 2T~
\~~ + r
2~ri h ~rnDl?r(t",)~/2
cot~r(t") 
r162
"
(6.11.12) 6.11.2.2 TimeDependent Radial Harmonic Oscillator. [251,259,440] (r as (6.11.10), ~(t) = ~/at 2 + 2bt + c, r(t) and/2 as in the previous example)
in
r(t")=r"
r(t')=r'
k"
{ 2 ~; im tr,,
(r'r" ~ 1/2
=
• exp
2=11~~,~i~+ ~;~]
,2 ~"~
]
) a/2
hsin I2r(t") /
&\it~'r
~ " (6.11.13)
6.11.2.3 TimeDependent 5Function Perturbation. [259,440] We consider a timedependent 5function perturbation according to V ( z ) =  7 5 ( x ) / ~ ( t ) , ~(t) : ~/at 2 + 2bt + c, however with w'2 = a c  b2 = O. We have the path integral identity (note r(t) : t/r ~(t")=z"
x(t,)=~, im
J x
(2rcihr(t"))
(x,,2~" t exp ~ \ ~ 7 ;
~7)
325
6.11 Explicit TimeDependent Problems
m7
~[ m7 (Ix"l
+ ~~exp   ~  \ ff,p + x erfc
~)
2 i hr(t") \ r
+
im72 .,,.])j
+ ~'F h T r ( t ' )
)1}
(6.11.14)
6.11.2.4 Hard Wall Potential. [180,259] We consider the example of a halfline (HL), i.e., L(t) _< x < oo, with the boundary moving according to L(t) = Lor The result then has the form [wP= 0, r as in (6.11.10)] ~(t")=~"
f
V(HL,~C,))x(t)exp \~j,,
~(t,)=x' 
2L0
21rihr(t") exp ~~ t x  ~ i T  x ' 2
(6.11.15)
x 2 fo~176 (6.11.16)
6.11.2.5 Rigid Box with One Wall Moving. [213,259] We consider the example of the infinite well (IW) with one boundary fixed at x = 0, and the other moving according to L(t) = Lo~(t). The result then has the form [w' = 0, O3(z, r) denotes a Jazobi theta function, r as in (6.11.10)]
~(t")=x" ~(t')=~'
2Lo
[~\
~" r
x [e~(~'lr 7Ei"Ir , .h~(t,,)~)~ _ e~(\ ."lr + ~'1r , .h~(t,,)~)J']~ (6.11.17)
326
Table of Path Integrals (~,(,,)1/2 [ira( ~" x,2~'~] 2Lo exp ~~ x"2 
x .~rq E sin (Trn x.~"'~ (~rn~'~ ( rr2n~ \ Lo ( " ] sin \ Lo ( ' ] exp  i h2~or(t"))
(6.11.18)
6.11.2. 6 Moving 5Function Perturbation. [276,440] x(t")=x" / ~(t')=x'
Dx(t'exp{~ft;"[~x2F~l'(xvt')]dt Fim
=
exp [2ff(x
,,
 x') ~
]
m7 m 7 ,, i m7 2 ] • exp  ~([x  vt" I + Ix'  vt'l) +
K~TJ
x erfr [IV[E{,x. ~qff \,
_ vt,,l + lx, _ vt, l _ ~.yT) ]
(6.11.19)
 h2 exp  ~ Ix"  vt"l + Ix'  vt'l i mv
h (x'vt')+
+~
dk exp
9
~  ~ (x"  vt") + i m
( 72
)]
T~ ~ +v2 T
ik2hT~ { "~m ,] eik(~"x')
[ imv( ~_)]} exp ik(I x"  vt" I + Ix '  vt'l) + ~ z"  x ' 1 + i kh~2 m7
(6.11.2o) 6.11.2. 7 Time Dependent HydrogenLike System. [874] (z' = 2x' e~'t', etc.)
x(t")=x" / ~(t')=~,
:Dx(t)exp{h fti" [2 (k2+ a2~z2 azk)
g~je~t/2]dt}
6.12 Point Interactions
327
= ~ zV~z,,oxp (  ~ t' q+ t" ) o
9
2vc~ "2n +
,
2),~)1
• 2 ( ~h 'v  2i~g m) "~ ~( z' : " ) ~i ~ exp [  ~2hV ' l  2ra " 0).
6.12.5.4 CoulombPotential in Two Dimensions. [443] (x = ( z , y ) ~ lit ~) "('")=""
[~t" ,
x(t')=x'
= V
axtj
I
~ 2 ~ h ~vg~ ~,o
+
m ~~
)~r2(~I_~) a+2 ~
1 @m oo . + 2~
0
a
) ]
+ l n  h
+ 27
r(l+~~)
~ ~e"(*"*') (20! l1 (6.12.45)
6.12 Point Interactions
341
6.12.5.5 The Free Particle in Three Dimensions: Feynman Kernel. [15,820] x(t")=x"
f
( i m [t"it2dt)
vro,~
exp \ 2h J,,
x(t')=x'
1 I~x'llx"al
= KC~ x K(~
e2"~nu/"(u + la  x ' l + Ix" al)
+ ]a  x' I + Ix"  al, 0; T) du ,
= K(O)(x,,,x,;T) +
ihT
K(~
 x'l + Ix"  a l , 0 ; T ) , (6.12.46b)
mia  x'llx"  al = K(~
", x'; T) + ~V(a)(xt)~P(a) (x ") e iE(~
1
fo ~176 e~l"laulm(u  la  x ' l  Ix"  al)
x K(~
 [a  x'l  Ix"  ah 0; T ) d u ,
+ l a  x'ff~"  "1
K(~
(6.12.46a)
( m y; T) = \ ~ ]
,~3/2
exp
(
m ix_yl2) 2ihT
(6.12.46c) ,
(6.12.47)
for a > 0, a = 0 and a < 0, respectively, the boundstate wave function is
~P('~)(x) = i
ah~ e2~aalxal/m m
with energy E(a)
ix _ al
,
a2h 6
E (~) = 2rr z mS
(6.12.48)
(6.12.49)
6.12.5.6 The Free Particle in Three Dimensions: Green Function. [17,443]
(x = (~, u, z) e IRa) x(t")=x" x(t')=x'
~ l x h"  x ' l )
2~'h 2 Ix"  xq exp
m)
+ ~
2
1
Ix"allax'l
+ 2~mE m
(6.12.5o) Here G(~
0) = m/27rh21x] (x E IR a \{0}).
342
Table of Path Integrals
6.12.5.Z Harmonic Oscillator in Three Dimensions. [443] (x = (x, y, z) E [a[ , 5 = ~ / h a , G(~)(E) denotes the threedimensional Green function of the harmonic oscillator, cf. Sect. 6.2.2.7)
]R3, v = ~1 + Elhw, a =
x(t")=x"
~L ~176 dTeiET/h i
~I'~
r'" (,:_
x(t')=x' = G(~
", x'; E) + G(~)(x"' a; E)G(~)(a, x~; E) (~) F.',.(E) m F (  v ) {I[Du(5)D,(_5)_ 2(27r)312h 2
Fa,a(E) : a
(6.12.51a)
D~(a)Du(5)]
,_,_
+
D~(a)D~,(f)t2Dv(a)D~(a)ID~,(a)D~, (  ) ]
}. (6.12.51b)
The case a = 0 gives: =
F[ 89 +
EIt~)],,,
m2 F2[89
f mw 2"~ ,
I'mw ~ ~
 ElhoJ)]2(EIh~ 89 47r3~4rr r##
x
m o< 
~~TV
c[~(l + 3/2 
EI~)]
~r[ 89 h
F[1(1/2
E/t~)] ~

EI~)]
~.(~,, r162162
(6.12.52)
6.12.5.8 Coulomb Potential in Three Dimensions. [17,443] (a = [a[, 5 = ax/&8mE/h, G(C)(E) denotes the threedimensional Coulomb Green function, cf. Sect. 6.8.6, n = e 2 x /  m / 2 E / h )
x(t,)=x, G(C)(x ", a; E)G(C)(a, x'; E)
= G(C)(x",x'; E) +
r.(?2(E)
(6.12.53a)
6.12 Point Interactions
Va,a(E) 
343
m r ( 1 ,~) 27rh 3 
o~
8~/Zg~
' (2a)M:~(2a) ," j.(2a) M~,89189 [2 W~,89  W~,,!(2a)M.,.
x


(6.12.53b) The case a = 0 gives:
4rrlir'r" ' i. 2E ~ " m) .
(6.14.27) 6.14.2 The Regularized Coherent State Path Integral. [220,512,596604,603,625]
p(t")=p" q(tl~)=q u
lim2~re~T/2  ~~oo
f
:Dp~(p,q)exp[lft]" ( ~(pdql _ qdp)  H(p, q)) dt ] 1"
p(t')=p ~ q(t~)=q ~
(6.14.28) p(tU)=p '' q ( t ' ) = q tt
:D#v(P' q) = 2rrvT exp 
2vT
p(tl)=p ' q(tl)=q '
(6.14.29)
and H(p, q) may be any polynomial Hamiltonian. 6.14.3 Path Integral for Spin System. [124,486] (n is a unit vector on S2, A is defined via VnAA(n)= nsuch that f:dtAh= f:dt f~ drn.h• n) JDEn(s)5(,nm21)exp [~0 ~a (A(n)fitn.B)sds] = ~
e 'lBlm (6.14.30)
6.14.4 Spin Quantization PhaseSpace Model. [355,360,486,725] (x = Acos tg, and cos dj intermediate in time between ~j1 and ~j) N~oo n~~ y j = l
•
A:J+e#Bxj)]
(6.14.31,
358
Table of Path Integrals
6.14.5 S p i n in a M a g n e t i c F i e l d . [302]
tl(t")=ft" f :D12(t)exp[iS/oT(C~176188176176 n(t,)=tt,
['Otto
t
vqtt
Ot e_ i(~,_BT_~,)/2 ] 2s
cos  ~ cos ~ e i(~''BT~')/2 sin ~ sin ~
=
(6.14.32)
6.14.6 C o h e r e n t S t a t e P a t h I n t e g r a l o n F l a g M a n i f o l d . [592] z' (t")=z"" Vz(tlVz'(t) z(t,)=z'
• exp m l o g L i ( z " * , ~ . ( e ' ) ) + m l o g L 2 ( z " * , ~ . ( e ' ) ) + i 
1 [ z 1 ~1 •
L(~.*,z,z',~.)dt
q'~'2 %2 '~
+ z 3 z3e
 zl z3 )~z2  zlz3)
}
, (6.14.33)
LI ( z " ' , z(t"))  (1 + Zltt*Zltt _~ Z2t,*Z2tt)
(6.14.34)
L 2(z ' " * , z ( t " ) )  " (1 + z 3"* z "3 + ( z 2"*  Z l "* %"*'" " )(z2"  zl" z3) (6.14.35) Z ~ i l + Z*" 2z2 + i n Z a*Z 3 "~ (Z~  Z*l Z 3*) ( Z 2' = lZ.__.337 ZlZ3) n ( z * , z) ~= i 77"t"1 "4[Zl ]2 AtIz2l 2 1 + ~3~E  zlz3p  (wxQl(z) + w2Q2(z)) , Izll 2 + I~l ~ Iz~  z l z 3 ? + n 1 + I z ~ p + Iz2[ 2 a+lz312+lz2z~z~l 2 , [~l 2 [z2[ 2 Q2(z) = ~ . 1 + [zll 2 + [z212 1 + [z3[~ + [ z 2  zlz3[ 2 " Ql(z) = m
(6.14.36) (6.14.37) (6.14.38)
6.14.7 C o h e r e n t S t a t e P a t h I n t e g r a l f o r SU(n). [360,733] ({p} are real constants, the Hamiltonian _H is defined via isospin functions Q m j is the magnitude of the (classical) isospin, and ~ are coordinates on the complex projective space IM = C P ( N ) with ]~12 = ~ =N1 ~ , ~ = C~)
6.14 Coherent States
359
C(t")=C"
f
~(t)~*(t)
~(t,)=~,
=
xexp
2J log(1 + ~"*~(t")) + i
1+
,c* "~' m  m e  i ~ T
i 1 + 1~12 /ira
exp
iJ
rn~l
"rn'm
1)/2T
'
(6.14.39)
rnI N
H(~*,~) = ~ / ~ , ~ Q m ( ~ . , ~ )
,
(6.14.40)
m=l rn1
2J Qm(~*'l~) = v / 2 m ( m 4 1) Z
k,~O
.m
. (ukuk

I mu~num) o=0+1~1%1/2 ' u m =uo~ra
[
V / 2 m ( m + 1) (m + 1)~km 4 O(k  m)
~ok =
]
(6.14.41) (6.14.42)
.
m1
6.14.8 G e n e r a l i z e d C o h e r e n t S t a t e s for SU(2). [298,619] (a(t), b(t) are determined via a =  i Aa 4 i fb*, b* =  i Ab  i fa* with the boundary conditions a(0) = 1, b(0) = 0) z ~(T)=z~
f z(O)=z2
Dz:Dz* (1 + z ~ z ( T ) ) J ( 1 + z*(O)z2) J 2J+1 2~ri (14 Iz12)= (14lZll2)(141z=l 2)
xexp (a*(T) 
J
~:zT~z d t  i
)
H(z*,z)ds
b*(T)z2 4 b(T)z; 4 a ( T ) z I. z~) 2J
(6.14.43)
(1 + Iz~l~)J(1 + Iz~l~)J
H(z*,z)  2J  A ( T ) ( 1  [z]2) 4 f ( t ) z * + f * ( t ) z
(6.14.44)
1 + Izl~
6.14.9 G e n e r a l i z e d C o h e r e n t S t a t e s for SU(1, 1). [298,383,619] z ~ (T)z
z(O)=z2
2 k  1 :Pz~z* (1  z~z(T))k(1  z*(O)z2) k 2~'i (1]z]2) 2 (1Iz11~)(1Iz212)
x exp
k
T 2"*Z Z*idt _ i 1 "z*z 

/o9
H(z*, z) ds
)
360
Table of Path Integrals
(1 Izxl2)k(1]z212) k
(6.14.45)
(a*(T) + b*(T)z2  b(T)z~  a(T)ZlZ2
H = 2A(t)K_ 0 + f(QK_+ + f * ( t ) K
2k
.
(6.14.46)
Here a ( t ) , b(t) are determined via a =  i A a  i f b * , b* =  i A b  i f a* with the boundary conditions a(0) = 1, b(0) = 0. K0, K+ span the ~u(1, 1) algebra [K_o, K_+] = +K_+, [K_+, K__] = 2K_ o.
6.14.10 C o h e r e n t S t a t e P a t h I n t e g r a l for A n y o n s . [475] (With the transformation ~ = x / ~ / V / 1 I(1~; and p describes via 0/Tr = 2#  1/2 the statistical behaviour, where 0 = 0 corresponds to bosons and 9 = ~r to fermions, respectively; I(I < 1) (* (T)=("
~
2/~1
:Df'(t):Df(t)
~(1 ICI2)
r162 xexp
1+1CI2)dt ] = e#'(T)e~i('+"~T~(0)
[~oT( ~*r p \ ii~~
2i~i~
(6.14.47)
6.14.11 S u p e r c o h e r e n t S t a t e P a t h I n t e g r a l for Osp(ll2; IR). [155,157, 657,822] (K_+, K0, _F+ are the generators of osp(ll2; IR.) with [K_0,_K+] = K+, [K_0,K_] =  K , [K_a~, K_] = +2K_0, [K0, F+] = + F + / 2 with a sign ambiguity in, e.g., K_ ; r r 0, 6, X, )C are Grassmann variables; T is the Casimir index) (z., r
exp [ _ i ft"
"1
= u*(e',t')+v(t",t')z'*T~*(t",t')z"+~(t",t')z'*z"+
(
• (l+z"*z"

~xr162 ]
\ + z'*z'  71~)1r ] q'r
H_(t) = A(t)K_ o + f(t)K_ + + f* (t)K__ + 0(t)_F+  0(t)F_ .
'5o(t",t')z'* (6.14.48) (6.14.49)
6.15 F e r m i o n s
361
The coefficients u(t), v(t), X(t) are determined by solving the coupled equations
(A(t)u(t) + f(t)v*(t) rk ~(~2x(t)) ,
u(t) =  i
~)*(t) = i (f(t)u*(t) + 1a(t)v*(t) + O(~2x(t)) ,
(6.14.50)
x(t) = ~i (O(t)u(t)  O(t)v* (t) ) . 6.15 Fermions 6.15.1 T h e Fermionic P a t h Integral.
6.15.1.1 The General Fevmionic Path Integral Via Coherent States. [217,313, 675,686,734,855,883]
)a)x(t): path integral in halfspace x > a with Neumann boundary
426
List of Symbols
condition at x = a. x(t")=z" f Dl~