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LQO TECHNION—93—QCAP The Many Faces of Quantum Chaos CERN l.IBRFIF!IES» GENE/Fl ||l||\\l||l||||\|||l|\\||||||||lllllllllllllllll ASHER PERES PBEBIBSEI Department of Physics, Technion—Israel Institute of Technology, 32 000 Haifa, Israel
Abstract — Various approaches to quantum chaos are reviewed and compared. It is not difficult to generate quantum evolutions (unitary mappings of Hilbert space) which are chaotic. Some of the algorithms achieving this result are quite formal and apparently devoid of physical interest. In the more promising approaches, the following property stands out: a quantum system whose classical analog is chaotic displays hypersensi tivity to small perturbations of its Hamiltonian. The long range evolution of such a system is unpredictable in the presence of small, uncontrollable perturbations. This unpredictability is the hallmark of physical chaos.
INTRODUCTION
The study of classical chaos is a mature and well understood discipline, covering a wide array of physical phenomena. On the other hand, the very existence of quantum chaos was until recently a controversial subject Indeed, the notion of chaos refers to the dynamical behavior as t —> oo, and has a fundamentally asymptotic character. On the other hand, for large enough but still finite times, the correspondence principle, which has undeniable heuristic value when we want to compare classical and quantum systems, fails for any nonlinear system In classical Hamiltonian dynamics, chaos arises because orbits may be unstable and errors in the initial data grow exponentially In quantum mechanics, the situation seems to be quite different. The dynamical law is a unitary evolution, 1/xt = U $0. Starting from a slightly different initial state 1,/26 yields 1/Jl = Ud}6, with the same unitary operator U. It follows that the scalar product of the perturbed and unperturbed states is constant: (1/1,,1%} : (1/JO , 1%}. In other words, small imperfections in the preparation of the initial state d0 uat grow. This elementary argument, however, is not convincing: if, instead of considering isolated classical trajectories, we use Liouville’s equation which describes an ensemble of classical systems, it readily follows from Koopman’s theorem [4] that the overlap of two different Liouville densities is constant in time, just as the overlap of two quantum wave functions, which also describe statistical ensembles. The mere constancy of this overlap therefore does not guarantee the absence of chaos. OCR Output in time, just as the overlap of two quantum wave functions, which also describe statistical ensembles. The mere constancy of this overlap therefore does not guarantee the absence of chaos. There is indeed a very simple way of generating a quantum chaos which closely parallels any type of classical chaos Consider an autonomous dynamical system obeying the equations of motion
dmk/dt:V(a:1,...,xN),k k:1,...,N. (1) lf N 2 3, such a system may be chaotic. Irrespective of its physical nature, it is always formally possible to introduce a Hamiltonian,
H:EVk(xl,...,xN)pk, (2) where the pk are new variables, defined to be canonically conjugate to the :1:k. This Hamiltonian obviously gives Eq. (1) as the law of motion. (Note that the Lagrangian, L E Z pk ick ——H, is numerically equal to zero. This is a highly constrained canonical system.) Quantization may then proceed as usual by the introduction of a wave function $(:101, . . . ,:1:N) and the substitution pk —>—i7i3/Oxk. We then still have Eq. (1) as the Heisenberg equation of motion for the operators wk, and since the latter commute (and therefore can be simultaneously diagonalized) any chaos in the solution of the classical equations (1) will be reflected as chaos in the time evolution of the expectation values We thus see that there is no formal incompatibility between quantum theory and chaos. The only relevant question is whether chaos can be found in “natural” quantum systems (in particular, in those which are experimentally observable) just as we encounter chaos in ordinary planetary systems, or in fluid mechanics. For example, a classical model of the lithium atom is chaotic; l1ow is this fact reflected in the properties of real lithium atoms?
QUANTUM ASPECTS OF CLASSICAL CHAOS
A quantum state is not the analog of a point in the classical phase space. The classical analog of a quantum state is a Liouville probability density. lf two Liouville functions are initially concentrated around neighboring points, and have some overlap, that overlap remains constant in time, by virtue of Koopman’s theorem Each one of tl1e Liouville functions may become distorted beyond description, until all phase space appears thoroughly mixed when seen on a coarse scale; yet, the overlap of these functions remains constant. However, the interesting problem is not how a Liouville probability density, which was initially given, will later overlap with the tortuous domain covered by the time evolution of another, initially given Liouville density. The experimentally relevant question is how each one of these time dependent domains overlaps with a fixed domain of phase space. ln a classical chaotic system, the final probability density, seen on a coarse scale, is homogeneous and roughly independent of the initial conditions. This property is called mixing. OCR Output Turning now to quantum theory, we may inquire whether Wigner’s function [6]
WM, P) = (rh)`N / z>(q — rl q + r) €2"’""’” dr, (3) which is the quantum analog of a Liouville density, is also subject to mixing. The answer is negative: Wl (q, p) has a much smoother time evolution than a Liouville function In particular, it can never develop contorted substructures on scales smaller than 71. Therefore, Wigner’s function VV(q, p) does not possess the mixing property, as defined above. In general, it is found empirically, by numerical simulations, that quantum mechanics tends to suppress the appearance of chaos. Quantum wave packets may remain localized, even when classical orbits are strongly chaotic, because the breakup of KAM (Kolmogorov-Arnol’d—Moser) surfaces starts in limited regions, and the remnants of these surfaces effectively act as barriers to quantum wave packet motion while permitting extensive classical flow A similar phenomenon appea.rs in simple models where the Hamiltonian includes a time~dependent perturbation In these models, which may have a single degree of freedom, the physical system is prepared in a state involving only one, or at most a few energy levels of the unperturbed Hamiltonian. One then finds that the time evolution of the quantum system involves only a few more neighboring energy levels, so that the energy remains "localized” in a narrow domain, even though no vestige remains of the KAM manifolds, the corresponding classical evolution is chaotic, and the classical energy increases without bound in a diffusive way [10]. As a consequence of this energy "localization,” the quantum motion is almost periodic and the initial state recurs repeatedly [11], as it would for a ti1ne—independent Hamiltonian with a point spectrum. The peculiarity here is not the recurrence itself, which is similar to that in a Poincaré cycle [12], and is completed only after an inordinately long time, but the fact that the quantum state after an arbitrarily long time can be computed accurately with a finite amount of work. This suggests a curious paradox [13]. Rather than computing a classically chaotic orbit by numeri cal integration of Hamilton’s equations, we could quantize in the standard way the classical Hamiltonian (with an arbitrarily low, but finite value of li) and then integrate the Schrodinger equation in order to follow the motion of a small wave packet. For example, we could integrate the evolution of the solar system for trillions of years, by assuming that the Sun and the planets, and all their moons, and all the asteroids, are point particles with constant and exactly known masses, and by replacing these classical points by Gaussian wave packets of optimum size. Could this be a less complex task (for t —+ oo) than the direct integration of Hamilton’s equations? Unfortunately, nothing can be gained by this subterfuge, because the initial quantum state from which we obtain the final wave packet is not itself a small wave packet, but is likely to be spread throughout all the accessible phase space. lf we want a genuine quantum simulation of a classical orbit of total duration t, we must start from a wave packet of size Aq ~ e’L‘(Aq)nna; and Ap ~ e“L’(Ap)nm;, where L is the Lyapunov exponent. Then, for a given value of the final uncertainty, the initial Aq Ap OCR Output must behave as exp(—2Lt), and this requires using a fake h which decreases as exp(—2Lt). As a consequence, we must take into account an increasing number of states, because the smaller Ti, the larger the density of energy levels for a given energy. A detailed analysis [13] then shows that the number of terms needed in the expansion into energy eigenstates increases exponentially with t, with at least the same Lyapunov exponent as for the classical problem. We thus see that if we attach to the word “chaos” the meaning that the computational complexity of a dynamical evolution increases faster than the actual duration of the motion (so that long range predictions are impossible, except statistically), then a genuine quantum system, with fixed Ti and a discrete energy spectrum, is never chaotic. On the other hand, if we want the correspondence principle to hold for a classically chaotic system, all we have to do is to use a fake value of h which decreases as e`2L’, where t is the total duration of the motion. It is crucial to specify which limit is taken first, t -—> oo or Ti —> 0 [14]. For fixed Fi, the correspondence principle is expected to break down after a finite time, of the order of log(S/Ti)/L, where S S: fpdq of the classical orbit.
UNITARY MAPS
An explicit model in which genuine quantum chaos occurs with properties similar to those of classical chaos is the configurational quantum cat map [15]. A charged particle moves in a square in the my plane, with periodic boundary conditions. It is subjected to time—periodic electromagnetic kicks (the electric field vector lies in the my plane, and the magnetic one is perpendicular to it). lt is found that the quasi-energy spectrum is absolutely continuous, and every value is countably infinite degenerate. The expectation value of the energy grows at an exponential rate, and a wave packet, initially concentrated in a small region of configuration space, gets exponentially stretched and folded, and is quickly distributed over the entire square. lt is even possible to have quantum chaos in a finite dimensional Hilbert space. Consider a unitary evolution, 1/2(t) = U(t) ¢, where the elements of the unitary matrix U(t) have a chaotic evolution, and spin j, and the three functions cz(t) are chaotic, like the functions mk(t) generated by Eq. Note that U(()) : ei°‘(°)‘J, and that i/¤(0) = U(()) 45. In this case, the meaning of “chaos” is that, for any given qi, two slightly different U(()) will evolve into widely different U(t), and thus yield very different state vectors Instead of chaotic parameters, we may also consider quasi—periodic ones. For example, it can be shown analytically [16] that a two-level quantum system, driven by a Fibonacci-like quasi-periodic perturbation, generically does not have a quasi—periodic time evolution. Rather, its evolution appears intermediate between quasi—periodic and chaotic, and it is characterised by fuzzy Poincaré sections, a slowly growing polarization fluctuation, and a singular continuous quasi-energy spectrum. Likewise, OCR Output numerical investigations of a kicked rotor model with three incommensurate frequencies demonstrate a transition from states which are localized in energy to extended states, at a critical value of the perturbation parameter [17]. The classical analog of these unitary maps is an ordinary rotation, generated by
r = w(t) >< r, (4)
where the angular velocity w(t) is a chaotic (or possibly quasi—periodic) function of time. This equation can be formally integrated, giving
1 _ r(t)=r0+i¥(a > The finite rotation angle a(t) is related to the angular velocity w(t) by [18] . . . a:a+-%?.(aXa)+&%9¥[aX(aXa)].I —‘‘ — HYPERSENSITIVITY TO PERTURBATIONS The most interesting approach to [quantum chaos is related to the following property: we are usually unable to perfectly isolate a microscopic quantum system from the influence of its macroscopic environment. In other words, the true Hamiltonian H is not known. The only thing we know is an idealized H0, which differs from H by a small perturbation, V, due to the imperfectly controlled environment. In that case, it is plausible that V has little influence on the quantum motion, 2/rg —+ tht, if the analogous classical system is regular; on the other hand, if the analogous classical system is chaotic, the final state 1/1, is likely to be strongly influenced by a small V. The state after a long time will therefore be unpredictable if V is unknown. This plausible scenario is supported by semiquantitative arguments which predict that perturbation theory fails—it does not converge—for quantum systems whose classical analog is chaotic [19]. The mechanism for this failure is the following: in a perturbation expansion, each successive step involves one more power of V, and one extra sum over states. For instance, the third order perturbation of an energy level is [20] V V it Vt I I nm m ri ’ [Vim E(3) : n g; g:Y i(Em _ — En) M if (Em " Eri)2 The essential difference between regular and chaotic systems is that the former have selection rules (corresponding to the classical isolating constants of motion) so that most matrix elements Vmn vanish (that is, they vanish for a reasonable V, having a classical limit A chaotic system, on the other hand, has no selection rules and the V,,,,, look like random numbers. A sum over states is OCR Output therefore like a random walk, and the ratio of consecutive steps in the perturbation series will be about (N<[Vm,,[2])l/2/E, where N is the number of terms to be summed, and is of the order of the dimensionality of Hilbert space. In this estimate, the symbol (|Vm,,|2} stands for the mean square of the various elements Vmn, and E is the typical energy scale of the system. For finite N, convergence of the perturbation series imposes a bound of order E / \/N for the "typical” matrix element Vmn [21, 22]. In the semiclassical limit (N—>oo) the perturbation series always diverge.; for any finite V. A simple model which has, in its classical version, both regular and chaotic domains, and which is described, in its quantum version, by a finite-dimensional Hilbert space, is the kicked top [23, 24]. As expected, if an initial state is concentrated in a classical regular domain, its dynamical evolution is not appreciably affected by a small perturbation of the Hamiltonian [21]. On the other hand, if the initial state is localized in the chaotic region, even near a fixed point, it quickly spreads throughout the entire chaotic domain [25], and its evolution is highly sensitive to perturbations [21]. Another interesting model is the quantum analog of the baker’s map [26]. In that case, it is possible to estimate the cost, from the point of view of information theory, of following the perturbed motion of a state vector in Hilbert space. It is found that the free—energy cost of tracking the perturbed pattern in fine—grained detail is enormously greater than the cost of the entropy increase that results from coarse graining. This marked discrepancy reflects the hypersensitivity to perturbations of a chaotic system. HEISENBERG PICTURE The Heisenberg picture, which uses a fixed state vector and time—dependent operators, is conceptu ally closer to classical Hamiltonian dynamics than the Schrodinger picture, with its fixed operators and time-dependent wave functions (the latter is -closer in spirit to classical statistical mechanics, with its time—dependent Liouville density). Let us therefore examine how chaos would appear in the Heisenberg matrix formulation of quantum mechanics. For example, a system with Hamiltonian H Z (P12 +1922 + $12 $22)/2» (8) has equations of motion $1 Z Pi (J = 1,2). (9) 2 iZ·¤¤j¤>k (k#J). both in classical and quantum mechanics. If as j and pj are classical variables, the orbits corresponding to these equations are chaotic for nearly all initial values. On the other hand, if xj and pj are operators satisfying [wm,p,,] : ihamn, (10) OCR Output the equations of motion have the formal solution Mft) = Ul¤>j(U)U» (11) P10) Z U[Pr(U)U The unitary operator U is explicitly given by U = §;exp(—iE,,t/li) Pu , (12) where the Eu are the eigenvalues of H, and the P,, are the projection operators on the corresponding eigenvectors. This explicit solution of the equations of motion in operator form is analogous to obtaining an ex plicit solution of the classical equations of motion for all initial conditions (while solving the Schrodinger equation for a given initial 1/1 would be like solving the classical Liouville equation for a given initial Liouville density). Where is the classical chaos hiding in this solution? The point is that in order to obtain explicitly the eigenvalues Eu and the corresponding operators Pu, we must solve an infinite set of equations, which is clearly impossible with finite computing resources. lf, as is usually done, we truncate the Hilbert space so as to obtain a large, but finite, number of equations, the commutation relations (10) are no longer valid. Alternatively, if instead of using the explicit, but formal, solution (11), we attempt to directly integrate the operator equations of motion (9), we find that the truncated equations are chaotic, even though the complete, exact equations are regular. Chaos thus becomes apparent if we use different truncation sizes: the result fluctuates erratically as we use matrices of larger and larger order. ODDS AND ENDS ln this review, the emphasis was on dynamical properties of chaotic quantum systems, and l did not mention the enormous body of knowledge on their static properties, such as the distribution of energy (or quasi—energy) eigenvalues [27]. It is known that the statistical features of these distributions may ultimately affect some dynamical properties [28]. However, there is one more type of quasi-quantum chaos that I wish to mention. There have been many attempts to generalize Schrodingefs equation so as to make it nonlinear. Sometimes, this nonlinearity involves a fundamental modification of quantum theory, and it may lead to considerable conceptual difficulties [29]. In other cases, the nonlinear Schrodinger equation only is an approximation: a many body system is described by an effective potential acting on each one of its constituents. Whatever the origin of nonlinearity, a nonlinear wave equation may produce ergodicity and positive Lyapunov exponents [30, 31] and in tha.t case too the result ma.y be called quantum chaos. OCR Output Acknowledgments —I am grateful to Shmuel Fishman and Joe Ford for enlightening comments. This work was supported by the Gerard Swope Fund and by the Fund for Encouragement of Research. REFERENCES J. Ford, Quantum chaos: is there any? in Directions in Chaos, edited by Hao Bai·Lin, World Scientific, Singapore (1987). A. Peres, Quantum Theory: Concepts and Methods, Kluwer, Dordrecht (1993) Chapt. 10. J. Ford and G. H. Lunsford, Stochastic behavior of resonant nearly linear oscillator systems in the limit of zero nonlinear coupling, Phys. Rev. A 1, 59 (1970). B. O. Koopman, Hamiltonian systems and transformations in Hilbert spaces, Proc. Nat. Acad. Sc. 17, 315 (1931). B. V. Chirikov, F. M. Izrailev, and D. L. Shepelyansky, Quantum chaos: localization vs. ergodicity Physica D 33, 77 (1988). 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