26th September 2002

An overview is provided over the physics of dielectric microcavities with non-paraxial mode structure; examples are microdroplets and edge-emitting semiconductor microlasers. Particular attention is given to cavities in which two spatial degrees of freedom are coupled via the boundary geometry. This generally necessitates numerical computations to obtain the electromagnetic cavity fields, and hence intuitive understanding becomes difficult. However, as in paraxial optics, the ray picture shows explanatory and predictive strength that can guide the design of microcavities. To understand the ray-wave connection in such asymmetric resonant cavities, methods from chaotic dynamics are required. |

1 Introduction

2 Dielectric microcavities as high-quality resonators

3 Whispering-gallery modes

4 Scattering resonances and quasibound states

5 Cavity ring-down and light emission

6 Wigner delay time and the density of states

7 Lifetime versus linewidth in experiments

8 How many modes does a cavity support?

9 Cavities without chaos

10 Chaotic cavities

11 Phase space representation with Poincaré sections

12 Uncertainty principle

13 Husimi projection

14 Constructive interference with chaotic rays

15 Chaotic whispering-gallery modes

16 Dynamical eclipsing

17 Conclusions

2 Dielectric microcavities as high-quality resonators

3 Whispering-gallery modes

4 Scattering resonances and quasibound states

5 Cavity ring-down and light emission

6 Wigner delay time and the density of states

7 Lifetime versus linewidth in experiments

8 How many modes does a cavity support?

9 Cavities without chaos

10 Chaotic cavities

11 Phase space representation with Poincaré sections

12 Uncertainty principle

13 Husimi projection

14 Constructive interference with chaotic rays

15 Chaotic whispering-gallery modes

16 Dynamical eclipsing

17 Conclusions

Maxwell’s equations of electrodynamics exemplify how the beauty of a theory is captured in the formal simplicity of its fundamental equations. Precisely for this reason, they also illustrate that physical insight cannot be gleaned from the defining equations of a theory per se unless we understand how these equations are solved in practice. In optics, one powerful approach to this challenge is the short-wavelength approximation leading to the ray picture. Rays are the solutions to Fermat’s variational principle, which in particular implies the laws of reflection and refraction at dielectric interfaces. As soon as one makes the transition from wave physics to this classical domain, concepts such as “trajectory”, “phase space” and “diffusion” become meaningful. In this chapter, special attention will be devoted to the classical phenomenon of chaos in the ray dynamics of small optical cavities, i.e., a sensitive dependence of the ray traces on initial conditions. The main implication for the corresponding wave equation is that it cannot be reduced to a set of mutually independent ordinary differential equations, e.g., by separation of variables. Even when only a few degrees of freedom are present in the system, their coupling then makes it impossible to label the wave solutions by a complete set of “quantum numbers” (e.g., longitudinal and transverse mode numbers). Such a problem is called non-integrable; this classification is due to Poincaré and applies both to rays and waves [1]. Integrable systems are characterized by a complete set of “good quantum numbers”, as will be illustrated in section 9

Microcavities can be broadly categorized into active (light emitting) and passive systems. In active devices, such as lasers, radiation is generated within the cavity medium, without any incident radiation at the same frequency; the cavity provides feedback and (in combination with a gain medium) sets the emission wavelength [2]. Some of the important properties that characterize a light emitting cavity are emission spectra, emission directivity, output energy, noise properties and lasing thresholds. Examples for passive systems are filters, multiplexers and other wavelength-selective optical components; their functionality relies on coupling to waveguides or freely propagating waves in their vicinity. To characterize and operate passive cavities, one is interested in transmission and reflection coefficients, i.e., in the resonator as a scatterer.

The two basic types of experiment, emission and scattering, for a given cavity geometry are intimately connected because they both probe its mode structure. The concept of modes in microcavities will be made more precise in section 8; as a working definition, let us denote as modes all electromagnetic excitations of the cavity which show up as peak structure in scattering or emission spectra and are caused by constructive interference in the resonator.

Nonintegrable cavities introduce added freedom into the design of novel optical components, especially when we apply results from the field of quantum chaos, in which modern quasiclassical methods are of central importance [3, 4, 5, 6, 7, 8, 9]. The term “quantum” in “quantum chaos” relates to the fact that cavity modes are discrete as a consequence of the constructive interference requirement mentioned above; the term “chaos” refers to a degree of complexity in the ray picture which renders inapplicable all simple quantization schemes such as the paraxial method.

The optics problem we are addressing is not one of “quantum optics”, but of the quantized (i.e., discrete) states of classical electrodynamics in spatially confined media. By quasiclassics, we therefore mean the short-wavelength treatment of the classical electromagnetic field. Confusion should be avoided between quasiclassics as defined here, and the semiclassical treatment of light in the matter-field interaction, which couples quantum particles to the electromagnetic field via the classical (vector) potentials: how we obtain the modes of a cavity (e.g., quasiclassically), should be distinguished from how we use them (e.g., as a basis in quantum optical calculations). Our main emphasis in this chapter will be on the “how” part of the problem, but the wealth of physics contained in these modes themselves points to novel applications as well. Having made this clarification, we henceforth use the term “semiclassics” synonymously with “quasiclassics” as defined above.

Many of the microcavities which are the subject of this chapter employ “mirrors” of a simple but efficient type: totally reflecting abrupt dielectric interfaces between a dielectric body and a surrounding lower-index medium. Resonators occurring in nature, such as droplets or microcrystals, use this mechanism to trap light. This allows us to draw parallels between nature and a variety of technologically relevant resonator designs that are based on the same confinement principle. A recurring theme in numerous systems is the combination of total internal reflection (TIR) with a special type of internal trajectory that skips along the boundary close to grazing incidence: the “whispering-gallery” (WG) phenomenon, named after an acoustic analogue in which sound propagates close to the curved walls of a circular hall without being audible in its center [10, 11]. An example for the use of the WG effect in cavity ring-down spectroscopy is reported in Ref. [12]. A WG cavity can provide the low loss needed to reduce noise and improve resolution in the detection of trace species; the trace chemicals are located outside the cavity and couple to the internal field by frustrated total internal reflection, or evanescent fields. In this section, we review Fresnel’s formulas before introducing the WG modes. This provides the basis for sections 9 - 16

In the ray picture, the Fresnel reflectivity, R_{Fresnel}, of a dielectric interface depends on the
angle of incidence (which we measure with respect to the surface normal) and relative
refractive index n. In particular, R_{Fresnel} drops significantly below the critical angle for total
internal reflection,

| (1) |

| (2) |

| (3) |

At higher orders of , wavelength-dependent corrections to Fresnel’s formulas do arise because TIR is truly total only for a plane wave incident on an infinite and flat interface. The latter does not hold for boundaries with finite curvature or even sharp corners, and similarly in cases where the incident beam has curved wavefronts, as in a Gaussian beam [13, 14, 15, 16]. The physical reason for radiation leakage in these situations is that light penetrates the dielectric interface to some distance which depends exponentially on , as in quantum-mechanical tunneling. This allows coupling to the radiation field outside the cavity [13, 14].

The analogy to quantum mechanics rests on the similarity between the Schrödinger and the Helmholtz equation for the field ,

| (4) |

One of the advantages of reducing the cavity problem to two degrees of freedom is that polarizations can approximately be decoupled into TE and TM. The resulting wave equations are then scalar, with the polarization information residing in the continuity conditions imposed on the fields at dielectric interfaces. Therefore, the fields are formally obtained as solutions of Eq. (4) in two dimensions. This scalar problem is the starting point for our analysis. In the absence of absorption or amplification, the index n(r) which defines our microcavity is real-valued, and approaches a constant (taken here as n = 1 for air) outside some finite three-dimensional domain corresponding to the cavity.

Figure 1 illustrates the formation of whispering-gallery modes (WGMs) for a dielectric ellipse illuminated by a TM polarized plane wave. The wavenumber k is made dimensionless by multiplying with the mean radius R of the two-dimensional shape. The sharp features in the spectra of Fig. 1 are caused by modes which rely on Fresnel reflection. This can be deduced from the fact that the spectrum becomes more crowded as the refractive index is doubled from n = 1.5 (a) to n = 3 (b); the intuitive reasoning is that a larger set of ray paths acquires high reflectivity as n increases. The more regularly spaced peaks of Fig. 1 (a) can in fact be identified as WGMs: Light circumnavigates the perimeter of the cavity in such a way as to insure TIR during a complete round-trip, i.e., χ>χ

The Mie treatment of circular and spherical micro-objects may serve as starting points for perturbative treatments of resonators whose shape deviates only slightly from rotational symmetry [35], or when objects in the vicinity of the cavity exert a weak influence [36]. Perturbation methods are among the most powerful tools of wave physics, but one must be aware that there are phenomena outside their reach (most apparent when energy denominators diverge). As a simple example, we return to the elliptical resonator [37]. If we assume impenetrable boundaries (of Dirichlet type), it is a classic textbook problem [38] to obtain the WGMs of the ellipse by applying perturbation theory to a circular cavity. The weakness of this approach is revealed when we recall that the modes can be found by an exact separation of variables and fall into two classes: WGMs and beam-like “bouncing-ball” states; in Fig. 2 we illustrate these two classes with slightly more complicated dielectric boundary conditions. Since bouncing-ball modes do not exist in the circle, perturbative expansions for this type of modes can be expected to become problematic, especially when attempting to describe how some modes of the circle lose their WG character under a continuous shape deformation.

The wave solutions shown in Fig. 2 show field intensity extending to the exterior of the cavity because the dielectric interface is “leaky”. In general, such a calculation must be performed numerically; we shall discuss the definition and emission properties of these leaky modes further below, in section 4 For now, we are only concerned with the internal intensity patterns. The reader will recognize a strong similarity between the “bouncing-ball” mode and a higher-order transverse Gauss-Hermite beam; this arises because the ray motion corresponding to this mode is a stable oscillation between the flat sides of the cavity. The analytic transverse form of the bouncing-ball beam in the ellipse is however not a Gaussian, but a Mathieu function. The caption of Fig. 2 gives the linewidth of the two types of modes, indicating that the WG resonance is almost an order of magnitude narrower than the bouncing-ball mode, despite the twofold higher refractive index used in Fig. 2 (b).

The ellipse has been chosen as an example because it is integrable in the limit of no leakage. Other oval deformations of the circle do not have this simplifying property. However, families of WG ray trajectories have been proven [39] to exist in any sufficiently smooth and oval enclosure, provided that its curvature is nowhere zero. This makes WGMs a very robust phenomenon of general convex oval cavities. Before we discuss in more detail the relation between the ray picture and the internal structure of the resonator modes, we now turn to some general considerations on what allows us to define the modes of an open cavity.

WGMs are not infinitely long-lived even in an ideal dielectric cavity, as we saw in the finite
linewidths of Fig. 1. In fact, for a finite, three-dimensionally confined dielectric body, all
solutions to Eq. (4) are extended to infinity, forming a continuous spectrum. A basis of
eigenfunctions is given by the scattering states, consisting of an incoming wave ψ_{in} that is
elastically scattered by the dielectric microstructure of index n(r) into an outgoing wave ψ_{out}. In
the asymptotic region where n = 1, the relation between incoming and outgoing waves
is mediated by the scattering operator S, known from quantum scattering theory
[40], to which the electromagnetic resonator problem is conceptually analogous.
The S-matrix formalism has long been in use in microwave technology as well as
optics [41]: A matrix representation is obtained by defining basis states |> (|>) in
which the asymptotic incoming (resp., outgoing) fields can be expanded; one has

Only the properties of the index profile, not of the particular incoming wave, enter S.
For a review of scattering theory see, e.g., [42]. As was seen in Fig. 1, the actual
cavity modes in this continuum of scattering states are revealed if we measure the
scattering of light as a function of wavenumber k. The amplitude of the scattered field
shows resonant structure at discrete values of k which do not depend on the detailed
spatial form of the exciting field ψ_{in}. These resonances are caused by poles of S in
the complex k plane, and are a characteristic of the microcavity itself. In Eq. (5),
we note that a pole of S admits nonvanishing ψ_{out} in the absence of any incoming
waves, ψ_{in} = 0. In these solutions at complex k, also known as the quasibound states,
we have finally found a proper definition of what we simply called “cavity modes”
earlier.

A well-known example is the linear two-mirror (Fabry-Perot) cavity [2]. Its resonances are
easily obtained within physical optics by writing the transmission of an incident ray as a
geometric series over multiple round trips. In each round trip, the amplitude of a ray in the
cavity accrues a factor r_{1}r_{2} exp(ink + i_{1} + i_{2}), where is the round-trip path length. We
have split the amplitude reflectivities at the individual reflections = 1, 2 into modulus r_{}
and phase _{}. Attenuation comes from r_{} < 1 ( = 1, 2 in the two-mirror case). The refractive
index n in the cavity is real, as stated below Eq. (4), and the wave number k is measured in
free space. The transmitted amplitude T_{total} is obtained by summing this over all repetitions,

| (6) |

| (7) |

| (8) |

| (9) |

| (10) |

We now discuss the significance of complex frequencies in Eqs. (4) and (5). This will highlight the relation between quasibound states and the observable cavity response to an external field or pump signal, which is of particular interest in spectroscopy. In Eq. (4), the wavenumber only appears in the form of a product nk, so that an imaginary part in ≡k + iκ≡/c can immediately be re-interpreted as part of a complex refractive index ñ at real k, setting

| (11) |

| (12) |

| (13) |

In spectroscopic applications, the quality Q and finesse F of a cavity are important figures of merit. Conventionally, one defines

| (14) |

| (15) |

| (16) |

| (17) |

| (18) |

| (19) |

| (20) |

| (21) |

| (22) |

| (23) |

| (24) |

| (25) |

| (26) |

The limit of a well-defined frequency in the wavepacket of Eq. (17) allowed us to assume
that the radiation interacts only with a single quasibound state, leading to the resonant delay
time Eq. (22). On the other hand, if we take |ξ(k)|≡C to be constant, the whole spectrum enters
with equal weight. One could still make the assumption that there is only a single isolated
resonance in the spectrum; then one immediately arrives at the well-known relation
between Lorentzian lineshape and exponential resonance decay: take the resonance
to be at wavenumber k_{r} (it is in fact always accompanied by a partner state [35]
at -k_{r}, but we can ignore it at large enough frequencies); replacing exp(i) = S
in Eq. (17) by Eq. (21), what remains is the Fourier transform of a Lorentzian,

| (27) |

The two limiting cases of (k) in the wavepacket Eq. (17) are just extremes of the time-frequency uncertainty relation. Experiments on microcavities have been performed both in the spectral and time domain. However, there has been only one report of the temporal nature of an ultra-short optical wavepacket (100 fs pulse, 30 μm short in air) incident on a pendant-shaped, hanging droplet with an equatorial radius of 520 μm [74]. Such pulses are spectrally broad but have the intriguing property that their spatial length is much shorter than the cavity size. Conceptually, the analogy to a well-defined particle trajectory suggests itself, and hence one looks for ballistic propagation in the cavity. In the experiment, wavepackets were observed to circulate inside the droplet in the region where the the WGMs reside; such paths are characteristic of high-order rainbows in that entry and exit of the light are separated by a large number of internal reflections [19, 47].

The time-resolved measurement showed that coherent excitation of a large number of
WGMs allows propagation of a short wavepacket along the sphere’s equator for several round
trips without significant decoherence, except for decrease in the wavepacket intensity because
of leakage of the WGMs. The novel aspect of the experiment that enabled the time-resolved
measurements was the use of two-color two-photon-excited Coumarin 510 dye molecules
embedded in the liquid (ethylene glycol) that formed the pendant droplet. The Coumarin
fluorescence (near 510 nm) appears when one wavepacket of _{1} and another wavepacket of _{2}
spatially overlap. Thus, the Coumarin fluorescence acts as a correlator between
wavepackets.

This time-resolved experiment was designed to answer the following questions: (1) does the excitation wavepacket remain intact and ballistic after evanescent coupling with WGMs? (2) after a few round trips, would dispersion cause the wavepacket to broaden? and (3) can the cavity ring down time be observed for those wavepackets that make several round trips? The answers were reached that the shape of excitation wavepacket remained intact, the wavepacket was not broadened after a few round trips, and that cavity ring down was observed for each round trip. This is an extension of the single-mode ring-down determined by Eq. (27).

In particular in the context of cavity ring-down spectroscopy, the usefulness of
time-domain measurements is recognized [75]. There, one deals with very high Q-factors and
their modification by the sample to be studied. Another application of temporal observation is
encountered in microdroplets, where the optical feedback provided by WGMs makes it
possible to reach the threshold for stimulated Raman scattering (SRS). The SRS
spectrum consists of sharp peaks, commensurate with the higher Q WGMs located
within the Raman gain profile. The highest Q value of the WGMs can be determined
either by resolving the narrowest linewidth of the SRS peak or by measuring
the longest exponential decay of the SRS signal after the pump laser pulse is off.
When Q > 10^{5}, the decay time = Q/ (where 3 × 10^{15} Hz for λ= 620 nm)
becomes a much easier quantity to measure directly because it is longer than 100 ps.
Otherwise, the spectral linewidth ( = /Q) needs to be resolved better than 0.006
nm (for = 600 nm). Cavity decay lifetimes as long as 6.5 ns have been observed
[76].

Having defined the density of modes for the open system in section 6, we can go one step further and perform the the average (k) over some finite spectral interval k. The suitable choice for such an averaging interval should of course contain many spectral peaks. Two relevant examples where an average spectral density is of use are the Thomas-Fermi model of the atom [77] and Planck’s radiation law. The latter shall serve as a motivation for some further discussion of the average (k) in the next section. For a presentation of the subtleties involved in this procedure, cf. Refs. [78, 79]. For our purposes, we simply remark that using Eq. (25) as a starting point, one way of arriving at an averaged spectral density is to make the formal substitution k k + iK. This amounts to an artificial broadening of all resonances as a function of k to make them overlap into a smoothed function; K then plays the role of the averaging interval [80].

In the limit of the closed cavity, the spectral density becomes a series of Dirac delta functions as the resonance poles move onto the real k axis and become truly bound states. The resonator then defines a Hermitian eigenvalue problem of the type we encounter in all electromagnetics textbooks, and many fundamental properties of realistic cavities can be understood within this lossless approximation. As a point in case, it is worth recalling the problem of blackbody radiation. From a historical point of view, this thermodynamic question was the nemesis of classical mechanics as the foundation of physics, because it led to the postulate of discrete atomic energy levels. From a practical point of view, the blackbody background can be a source of noise in spectroscopic measurements, and its spectrum is modified by the presence or absence of a cavity.

From an electrodynamic point of view, the central nontrivial aspect of Planck’s problem is that the average spectrum of the blackbody can be observed to be independent of the cavity shape. The explanation of this universality rests on the average spectral density of cavity modes, which is found to be independent of the resonator geometry to leading order in frequency. Although it may be intuitively convincing that the shape of the enclosure should become unimportant when its dimensions are large compared to the wavelength [81], the actual proof requires a large measure of ingenuity. In a series of works beginning in 1913 [82], Weyl showed that for a closed, three-dimensional electromagnetic resonator of volume V , the average spectral density as a function of wavenumber k is (including polarization)

| (28) |

| (29) |

Stability is a property of particular rays, not of a cavity as a whole. To understand all the modes of a generic cavity, one has to go beyond paraxiality. It also must be realized that paraxial optics is itself a special case of a more general approximation scheme, known as the parabolic-equation method [85, 87]. This name refers to the fact that the resulting wave equation of Schrödinger type is mathematically classified as a parabolic differential equation. This is mentioned here because at a more abstract level, one can build approximations of paraxial type not only around rectilinear rays: e.g., for WGMs, an envelope function ansatz in polar coordinates, Ψ(r, φ) =χ (r;φ ) exp(-iβφ) allows to become the “propagation direction”. The question of whether a mode can be called paraxial or not then becomes dependent on the coordinate system one uses (e.g., Cartesian vs. cylindrical). A more fundamental distinction by which the cavity as a whole can be classified is that of integrability. Next, we discuss some examples for this concept.

One feature that can be observed in both Fig. 2 (a) and (b) is that the modes exhibit
caustics inside the dielectric, i.e. well-defined curves of high intensity which in the
ray picture correspond to envelopes at which the rays are tangent. In the WGM,
the caustic is an ellipse confocal with the boundary; it can be parameterized by
its eccentricity, e_{c}. In Fig. 2 (b), the caustic consists of two confocal hyperbola
segments. Caustics separate the classically allowed from the forbidden regions, in the
sense of the WKB approximation. In this section, we discuss some examples of how
quasiclassical and exact solutions can be obtained in non-chaotic but nontrivial cavities, if
coupling to the exterior region is neglected. In this closed limit, all fields can be
written as real-valued functions obeying standard boundary conditions (Dirichlet or
Neumann).

Figure 3 (a) illustrates the WG caustics of an ellipse and their relation to the quasiclassical quantization method of Einstein, Brillouin and Keller (EBK, the multidimensional generalization of the WKB approximation) [88]: given a particular caustic of eccentricity e

| (30) |

| (31) |

Even in three-dimensional integrable cavities, the EBK method can yield highly accurate results down to the lowest-frequency modes. As an example, we mention recent work on a microlaser cavity with strong internal focusing properties, which consists of a dome in the shape of a paraboloid, on top of a layered semiconductor [90], cf. Fig. 3 (b). As in the ellipse, the exact solution for the mode shown for this plano-parabolic mirror geometry bears some resemblance to the more familiar paraxial optics, in this case a Gauss-Laguerre beam. However, just like the circular resonator, the parabolic dome has no stable ray orbits, owing to the confocal condition. The modes can be found exactly because the geometry allows separation of variables in parabolic cylinder coordinates.

The short-wavelength approximation can be made highly accurate (as in the parabolic dome) or even exact (as in the Fabry-Perot cavity of section 4). As a nontrivial generalization of exact quantization based on rays, Fig. 3 (c) shows the equilateral triangle. All its modes can be obtained by superimposing a finite number of suitably chosen plane waves [10, 91, 85]. This is achieved by “unfolding” the cavity into an infinite lattice created by its mirror images. These examples show that ray optics, as the skeleton which carries the wave fields, remains a useful tool far beyond the paraxial limit.

However, the reader may ask: what is the use for semiclassical methods in exactly solvable problems such as the above systems? After all, semiclassics is simple in separable systems! Beyond quantitative estimates, the value of quasiclassics is that the connection between modes and rays can be carried over to deformations of the cavity shape where the separability is destroyed. When this happens, eigenstates cannot be labeled uniquely by global quantum numbers anymore. However, as Weyl’s formula teaches us, the absence of good quantum numbers does not imply the absence of good modes. In order to classify the latter, we shall attempt to label them according to the ray trajectories to which they quasiclassically correspond.

Taken together, the discussions of stable resonators in the previous chapter [86] and of integrable systems in the foregoing section provide the essentials of “conventional” resonator physics. However, in the infinite space of possible cavity shapes, most geometries are nonintegrable and hence display chaos in their ray dynamics, as mentioned in the introduction to this chapter. To illustrate the transition to chaos, Fig. 4 shows the continuous evolution of an individual quasibound state as the shape of an oval resonator is deformed. The cavity shape is a two-dimensional quadrupole of mean radius R. Now the distinction between the mathematical ellipse and other oval shapes becomes important. Although the difference between the shapes in Figs. 4 and 2 is barely discernible at small , the quadrupole is not an integrable cavity and displays a far more intricate internal mode structure. In particular, caustics become frayed, and nodal lines form ever more complicated patterns as increases.

Given this complicated scenario, it is not immediately clear that ray considerations can help at all in understanding the properties of states such as those in Fig. 4. However, it turns out that quite the opposite is true: it is the added complexity of the internal ray dynamics that can be identified as the cause of the more complex wave fields. We make this somewhat provocative statement because the previous examples have proven the success of using ray considerations as a scaffolding for constructing the cavity modes. Unfortunately, there is so far no complete theoretical framework for the quasiclassical quantization of partially chaotic systems; the EBK method, which in quantum mechanics gives rise to the corrected Bohr-Sommerfeld quantization rules [88], fails when the system does not exhibit well-defined caustics, as was already noted by Einstein [92].

However, it is worth following the quasiclassical route because there exists a vast amount of knowledge about the classical part of the problem: chaotic ray dynamics can help gain insights into the wave solutions that cannot be gleaned from numerical computations alone. Ray optics can be formally mapped onto the classical mechanics of a point particle; this allows us to leverage a rich body of work on chaotic classical mechanics – a mature, though by no means complete field [1, 5, 93, 94]. Hence, quasiclassics in the presence of chaos is a challenging undertaking, but also a useful one because the numerical costs for obtaining exact wave solutions in nonintegrable systems are so high.

As observed in section 7, pulses that excite many cavity modes can behave ballistically, i.e. seem to follow well-defined trajectories. The idea of the quasiclassical approach is complementary to this: from the behavior of whole families of ray trajectories, we want to extract the properties of individual cavity modes.

As the ellipse already taught us, different types of ray motion can coexist in a single resonator – e.g., WG and bouncing-ball trajectories. In such cases, it is desirable to know what combination of initial conditions for a ray will result in which type of motion. Initial conditions can be specified by giving the position on the boundary at which a ray is launched, and the angle it forms with the surface normal. The proper choice of initial conditions was identified at the end of section 4 as a prerequisite in the quasi-classical modeling of resonance decay. The present section introduces the tools necessary for solving this problem.

Information on the possible types of ray motion, and how they are grouped in families, is contained in the Poincaré surface of section (SOS), cf. Fig. 5. It is a representation of the classical phase space in the - sin plane, spanning all possible angles of incidence and reflection positions along the boundary. The definition of in the center of Fig. 5 includes a sign, measured positive in the direction shown by the arrow from the outward normal. This reflects the observation made below Eq. (31) that sin is proportional to the z component of the instantaneous angular momentum at the reflection, and the sign thus distinguishes the sense of rotation.

By combining the two variables and sin , dynamical structure can be revealed which would remain hidden in a collection of real-space ray traces. Non-chaotic trajectories are confined by local or global conservation laws to one-dimensional lines, called invariant curves. Almost all rays in the ellipse follow such curves, as seen in Fig. 5 (a), which shows a clearcut division into oscillatory and rotational (WG) motion, cf. also Fig. 2. This is analogous to the phase space of a physical pendulum [95]. The separatrix between the two types of motion corresponds to a diametral ray orbit connecting the points φ= 0, π on the boundary.

Rays launched on an invariant curve must remain on it for all subsequent reflections. Such unbroken curves persist for | sin | 1 even in Fig. 5 (b); this is just the WG limit. A lesser degree of robustness is observed for the oscillatory trajectories: they are surrounded by elliptical “stable islands”, but a chaotic sea forms in-between, owing to the fact that in Fig. 5 (b) the diametral separatrix orbit mentioned above is unstable, developing the sensitivity to small deviations in initial conditions which is typical for chaotic behavior. Chaos develops preferentially around such separatrix orbits. The analogy to a pendulum makes this plausible: there, the separatrix is the unstable equilibrium point at which the pendulum balances upside down. Note that the different types of motion described here are mutually exclusive, i.e. chaotic orbits never cross over into the islands of stability. This has the important effect that chaotic motion is indirectly affected by the presence of stable structure in the SOS.

The central reason why the SOS is introduced in this chapter is that it allows us to form a
bridge between the discrete electromagnetic modes of the cavity and their measurable emission
characteristics. Emission means coupling to the environment and hence appears as a
dissipation mechanism in the internal cavity dynamics, as illustrated in Eq. (9). Recalling
section 2, the emission from a dielectric cavity is governed foremost by Fresnel’s formulas,
which give a wavelength-independent relation for the reflectivities r_{} along any ray path,
determined only by the local angle of incidence _{} at reflection ; in the example of the
Fabry-Perot cavity, nothing else is needed to determine the decay rate of a cavity mode,
cf. Eq. (8). Now we note that the SOS shows the angle of incidence on the vertical axis, and
hence the Fresnel coefficient of any given reflection can be read off with ease. In
particular, the critical angle for TIR, Eq. (1), is represented in the SOS as a horizontal
line. When this line is crossed from above (in absolute value), refractive escape
becomes possible. The SOS then tells us which combinations of initial conditions give
rise to ray trajectories that eventually reach this classical escape window in phase
space.

To complete the bridge between a cavity mode and its corresponding set of initial conditions in the ray dynamics, we have to find a way of projecting quasibound fields onto the SOS. To this end, let us briefly discuss some aspects of how mode fields can be measured. Individual quasibound states can be studied in great detail in microlaser experiments [96, 97, 16, 98], because one can make spatially and spectrally resolved images of the emitter under various observation angles. As is evident from Fig. 4, the far-field intensity depends on the polar angle of the detector relative to some fixed cross-sectional axis of the object (say, the horizontal axis). Instead of simply measuring this far-field intensity, however, one can also ask from which points on the cavity surface the collected light originated – i.e., an image of the emitter can be recreated with the help of a lens. The dielectric will then exhibit bright spots at surface locations whose distribution in the image may change as a function of observation angle . The two variables and are conjugate to each other, because measures a propagation direction whereas is a position coordinate of the cavity.

The conjugacy between and means that they are incompatible, in the sense that they obey an uncertainty relation: the field distribution as a function of in the image can be deduced from the far-field distribution in by a Fourier transformation, and that is the function performed by the imaging lens. But quantities related by Fourier transformation cannot simultaneously have arbitrarily sharp distributions. Physically, in an image-field measurement a large lens is needed to get good spatial () resolution, but this leaves a larger uncertainty about the direction in which the collected light was traveling [98].

If we could plot the measured intensity as a function of both and , we could generate a
two-dimensional distribution similar to the SOS of the preceding ray analysis. In fact, knowing
the surface shape, it is only a matter of trigonometry and the law of refraction to
transform the pair (, ) measured on the outside to (, ) inside the cavity and
hence make the analogy complete. But if one has already calculated the wave field of
a quasibound state, what is the advantage of representing it in this phase space
rather than in real space? The answer is that the real-space wave patterns often
obscure information about the correlations between the two conjugate variables
(, sin ). As mentioned above, the SOS which plots these variables would allow us to
understand whether a given mode is allowed to emit refractively; and if so, we can
furthermore determine at what positions on the surface and in which directions
the escape will preferentially occur. This is determined by the joint distribution of
(_{}, sin _{}) over all reflections encountered by the rays corresponding to the given mode
[51].

To begin a phase space analysis, we now return to the question of how to extract correlations between conjugate variables from a wave field. This can be illustrated by analogy with a musical score: any piece of music can be recorded by graphing the sound amplitude as a function of time. On the other hand, musical notation instead plots a sequence of sounds by simultaneously specifying their pitch and duration. These are conjugate variables because monochromatic sounds require infinite duration [99], but their joint distribution is what makes the melody. The reason why musical scores can be written unambiguously is that they apply to a short-wavelength regime in which the frequency smearing by finite “pulse” duration goes to zero (in this sense, all music is “classical”. . . ). In the same way that musical notes are adapted to our perception of music, phase-space representations of optical wave fields project complex spatial patterns onto classical variables relevant in measurements.

This reasoning is familiar from quantum optics as well: there, amplitude and phase of the electromagnetic field are conjugate variables, and their joint distribution is probed in correlation experiments [100, 101]. A possible way of obtaining a phase-space representation is the Husimi function, obtained by forming an overlap integral between the relevant quantum state and a coherent state corresponding to a minimum-uncertainty wavepacket in the space of photon number versus phase. Here, we want to extract the same type of information about the joint distribution of the conjugate variables , sin relevant to an optical imaging measurement on a cavity mode, as described above. By projecting the electromagnetic field onto the SOS, measurable correlations are revealed which can be compared to the classical phase space structure. This is another realization of a Husimi function; the examples mentioned above differ essentially only in the actual definition of the coherent state basis onto which the wave field is projected [102].

Motivated by the above remarks on measurement of emission locations versus detector
direction, we now note that the coherent states with which the cavity modes should be
overlapped are in fact the Gaussian beams [2]. The fundamental Gaussian beam ψ_{G} has the
property of being a minimum-uncertainty wavepacket in the coordinate x transverse to its
propagation direction z, evolving according to the paraxial (Fresnel) approximation in
complete analogy with a Schrödinger wavepacket,

| (32) |

| (33) |

Although individual Gaussian wavepackets are not good solutions of the wave
problem inside the cavity, they are always an allowed (though overcomplete) basis
in which the field can be expanded. This is all we require in order to obtain the
desired, smoothed phase space representation of ψ_{int}: The value of the overlap integral,

| (34) |

In this section, we return to the “mystery of the missing modes” with which section 8 confronted us: Weyl’s density of states is asymptotically the same for all cavities of the same volume, even if we cannot classify their modes according some numbering that follows from integrability, or at least paraxiality. Figure 4 shows that a given mode does not cease to exist as integrability is destroyed with increasing , but evolves smoothly into a new and complex field pattern with a well-defined spectral peak that can even become narrower with increasing chaos in the ray dynamics. As noted in section 10, the “mystery” is how at large shape deformation the increasingly numerous chaotic rays, which crisscross the cavity in a pseudo-random way and even diverge from each other, can give rise to constructive interference, and hence to a resonant mode.

In section 9, the quasiclassical quantization approach led from the ray picture to the wave fields. Now, we follow the opposite direction and use the Husimi projection of a given cavity mode to perform a “dequantization” which leads to the ray picture. In this way, the novel physics of chaotic modes can be described without much technical detail.

By superimposing the Husimi projection of a mode onto the SOS for the same deformation, we can identify classical structures on which the mode is built. We have so far encountered two types of structure, cf. the caption of Fig. 5: (I) invariant curves such as those formed by WG rays, permitting EBK quantization as discussed in section 9; (II) stable islands around periodic orbits, for which the paraxial approximation of the previous chapter [86] can be used. In both cases, a finite (and usually small) number of wavefronts suffices to achieve the constructive interference without which no cavity mode can form. The mode shown in Fig. 6, on the other hand, overlaps with neither of these non-chaotic phase-space components; it is a chaotic mode. One possible way to understand that chaotic modes are possible at all, is to realize that even the chaotic sea is not structureless.

The classical objects of crucial importance for an understanding of chaotic modes are the unstable periodic orbits of the cavity. They give rise to a third type of invariant sets in the ray dynamics: (III) stable and unstable manifolds. These are one-dimensional curves which, like the structures (I) and (II), have the property that rays launched anywhere on the manifold will always remain there. Along these manifolds, trajectories rapidly approach or depart from a given periodic orbit, such as the one shown in Fig. 6 (b, right inset). A non-periodic ray trajectory which moves along one of the corresponding manifolds is displayed in the top center inset.

The stable and unstable manifolds form an intricate web, the “homoclinic tangle”, which was already recognized by Poincaré as the cause of severe difficulties in calculating the dynamics [5]. Part of this tangle is shown in Fig. 6 (b), forming interweaving lobes which play an important role in controlling the transport of phase-space density. A discussion of this “turnstile” action and the relation to the stability of the orbits from which the tangle originates is found, e.g., in Refs. [1, 108]). As the Husimi projection in Fig. 6 (b) shows, the wave intensity is guided along the invariant manifolds of the unstable periodic orbit [110], imparting a high degree of anisotropy onto the internal intensity and on the emission directions.

Particularly strong Husimi intensity is found at the reflection points of the periodic orbit shown in the inset to Fig. 6 (b); this corresponds to the high-intensity ridges in the real-space intensity of Fig. 4 (d). Such wavefunction scarring [109] is a surprising feature, because a wave field concentrated near the origin of the homoclinic tangle, i.e., at the unstable periodic orbit itself, spatially appears similar to a sequence of Gaussian beams, despite the fact that paraxial optics requires the underlying modes to be stable. Even if we do not start from this paraxial point of view, but follow the historical developments in quantum chaos, scars seem to run counter to the early conventional wisdom that asserted chaotic modes should posses a random spatial distribution [108]. In a recent study on lasing in a large, planar laser cavity with no stable ray orbits [111], laser operation with focused emission was observed and explained by wave function scarring. This phenomenon is of interest not only from the applied point of view, but also because a complete understanding of the quasiclassical theory for modes corresponding to the invariant manifolds of type (III) is as yet missing [112, 113].

The lifetime of the quasibound states in Fig. 4 (d) yields a Q-factor of Q≡k/κ = 700, which is of the same order as that of a high-power microlaser [97] that was demonstrated recently as an application of cavity design using the concepts introduced in section 11. The bowtie-shaped mode encountered there was of paraxial type, centered around a stable periodic ray orbit. If low lasing thresholds and hence higher Q-factors are desired, we recall from Fig. 2 that WGMs are preferred. As shown in Fig. 7, WGMs in partially chaotic cavities display high Q and strongly anisotropic emission patterns. Comparing the non-chaotic case of Fig. 2 (a) and the chaotic cavity of Fig. 7, we indeed find stronger focusing in the latter case.

Chaotic WGMs are among the few types of chaotic modes for which an approximate quasiclassical description along the lines of section 9 is possible. They correspond to ray trajectories which are part of the chaotic sea but circulate along the cavity perimeter many times before exploring other phase-space regions, cf. Fig. 5 (b), bottom. There is thus a separation of time scales between the fast WG circulation and a slow deviation from WG behavior which inevitably ends in chaotic motion. This makes it possible to use an adiabatic approximation [114] in which the “diffusion” away from WG motion is neglected for times long enough to contain many round-trips of the ray.

As a consequence, one can formulate approximate quantization conditions for chaotic WG modes by ignoring the chaotic dynamics [51], leading to equations of the EBK type, Eq. (30). The approximate nature of this quantization can be recognized in Fig. 7 (a), which differs from the ellipse-shaped cavity of Fig. 2 (a) in the fact that a small but nonvanishing field persists in the cavity center. This indicates that there no longer is a well-defined caustic, as we required in section 9 Nevertheless, the Husimi distribution (grayscale) in Fig. 7 (c) condenses approximately onto a one-dimensional curve (the shape of which can be given analytically within the adiabatic approximation [51]), following the homoclinic tangle [the SOS in Fig. 7 (c) is vertically expanded compared to Fig. 6 (b)]. The location of this adiabatic invariant curve is determined by the EBK quantization; its minima are seen to approach the TIR condition.

As indicated in section 14, chaos does become important when the emission properties of a chaotic WG mode are concerned. It is the deviation of the rays from the above adiabatic assumption which allows an initially confined ray to violate the total-internal reflection condition after some time, and hence escape refractively, cf. Fig. 7 (b). In Fig. 7 (c), this escape results from wave intensity leaking across the TIR condition near φ= 0, π. This can be reproduced within a ray simulation by launching an ensemble of rays on the adiabatic invariant curve, and recording the distribution of classically escaping rays.

This implements the program outlined at the end of section 4 The stochastic branching of a ray trajectory into transmitted and reflected parts, also known as ray-splitting [48], could be summed up as a geometric series in the Fabry-Perot example, Eq. (6). In the oval cavity, Monte-Carlo simulations for large ensembles of incident rays are required [51]. This ray approach has been applied in scattering problems along the lines of Eq. (10) [49]; the value and applicability of ray considerations in scattering from chaotic optical cavities was demonstrated in Ref. [50]. The difference to the quasibound-state approach is that our incident rays are launched inside the cavity, not outside of it.

Because chaotic diffusion makes classical escape from WGMs possible in the first place, the ray approach in fact becomes more useful at large deformations where chaos increases. That is just the regime where exact scattering calculations encounter difficulties. Starting from the ray approach to emission directionality and decay rates, one can then successively introduce corrections at higher orders in . In particular, in Ref. [51], the openness of the system was taken into account in the semiclassical quantization step by introducing additional phase shifts into the EBK equations. The reason is that penetration into the outside region, as mentioned in section 2, creates an enlarged, effective cavity boundary which lengthens the ray path lengths. This “Goos-Haenchen effect” [15] is well-known from plane interfaces as well. The effective cavity is actually of slightly different shape than the physical geometry. This has the consequence that even the elliptic cavity of Fig. 2, which for impenetrable boundaries is integrable, becomes non-integrable as a dielectric body [46, 47].

In the 0 limit, the internal dynamics of the ellipse displays no chaos, cf. Fig. 5 (a). This makes it an important test case with which to compare whenever we want to identify fingerprints of chaos. The power of such comparisons has been illustrated in an experiment on prolate, dye-doped lasing microdroplets [98]. The geometry of the droplets was rotationally symmetric around the polar axis, which is also the major axis. If one analyzes only the light propagating in a plane containing the two poles, then the cavity is effectively formed by the 2D axial cross section whose shape is an oval of the type we have been discussing above, cf. Fig. 5. The experimental technique of recording with a CCD camera an image of the micro-object at various observation angles (measured with respect to the polar axis of a spheroidal droplet) reveals prominent bright and dark regions along the cavity rim, from which the position and exit angle of the laser emission can be extracted simultaneously.

As discussed in section 3, the lasing modes in low-index materials such as droplets should
be of WG-type. The experiment illustrates both the focusing property, as well as the
robustness of WGMs. When the lasing droplets are imaged from the side perpendicular to the
polar axis (θ= 90^{o}), one thus expects emission directionality of the type shown in Figs. 2 (a) or
7 (a), showing emission from the points of highest curvature (i.e., the poles). However, neither
case describes the observed emission correctly. The droplets appeared dim at the poles instead
of being brightest there.

This counter-intuitive phenomenon can be explained by the phase space structure shown in
Fig. 5 (b). An essential role is played by the small islands belonging to the stable
four-bounce diamond-shaped orbit depicted in the Figure. Its effect is that WG rays
of the type shown below the SOS are prevented from reaching the high-curvature
points (Φ= 0, π) at angle of incidence sinχ = 0.65, because chaotic and stable motion
are mutually exclusive (cf. section 11). Now _{TIR} = arcsin 0.65 just happens to
be the critical angle for the droplet-air interface; hence, near-critical escape from
the highest-curvature points is ruled out for chaotically moving WG rays in the
droplet. In Fig. 5 (b), it is impossible to see from the ray trace in real space that the
combination φ= 0, χ=χ _{TIR} will never occur. Likewise, this information is by no means
obvious from the wave equation itself Only the phase-space analysis using the SOS
reveals the small islands that cause this effect, which has been called dynamical
eclipsing. The emission profile is determined by the phase space structure in the vicinity
of the emission line sin _{TIR}, and therefore we do not detect dynamical eclipsing
in the higher-index example of Fig. 7. A glance at Fig. 5 (a) makes it clear that
the effect will never occur in the ellipse, because the relevant islands do not exist
there.

The subtle difference in shapes between ellipsoid and quadrupole raises the question: How short does the wavelength have to be compared to the cavity dimensions, in order to be able to resolve such differences in geometry? Clearly, the droplets are so large compared to the optical wavelength (35 m diameter for an equivalent-volume sphere) that we are deep in the short-wavelength limit. What the above example teaches us is that it is not the structure in real space that has to be resolved by the waves, but the structure in phase space. The uncertainty principle will limit the size of the islands that can have a noticeable effect on the cavity mode structure, but the important thing to realize is that small differences in cavity shape can lead to large differences in phase space structure. As a result, the phase space effects introduced in this chapter turn out to be observable even in cavities that are not much larger than the wavelength. This is an extension of the claim made for integrable systems in section 9, that quasiclassical methods often extend all the way to the longest-wavelength modes.

Future work on nonintegrable dielectric cavities can proceed in various directions: our understanding of the classical ray transport in partially chaotic systems has to be explored further, and the systematic wave corrections to these ray ideas must be investigated. Two such corrections which are of particular importance in chaotic systems are dynamical localization [114, 51, 115] and chaos-assisted tunneling [107]. Dynamical localization is an interference effect that suppresses decay rates below the value expected from ray simulations as outlined in this chapter. It becomes more pronounced in smaller cavities, and in fact helps maintain high Q-factors in the presence of chaotic ray dynamics. Localization arises in many areas of physics as a name for wave corrections to a classical diffusion picture. In our case, the diffusion happens in the chaotic phase space of the ray dynamics. Chaos-assisted tunneling, on the other hand, acts to enhance radiation losses of non-chaotic WGMs at small deformations when chaos pervades only the low-sin regions of the SOS. It also leads to a coupling between WG wavefronts encircling the cavity in counterclockwise and clockwise sense of rotation, e.g., between the extreme bottom and top of the SOS in Fig. 4 (a). In the ray picture, spontaneous reversal of rotation direction is forbidden for WG trajectories on invariant curves; the mechanism at work here is tunneling in phase space.

The phase space approach to microcavity electrodynamics provides a means of classifying
the wide variety of modes in cavities where “good quantum numbers” do not uniquely
enumerate the spectrum. Semiclassical methods provide the connection between
individual modes and manifolds in phase space, such as the caustics in section 9,
belonging to invariant curves in the SOS. Open questions remain about how to
establish this connection in the general case when caustics are broken up by ray chaos.
The particularly important case of WGMs, however, is amenable to approximate
treatments which permit quantitative predictions. Many systems in which strong
matter-field coupling enables quantum-electrodynamic studies, are actually large
enough to be in the quasiclassical regime: examples are semiconductor domes [90]
(where the effective wavelength is shortened by a high refractive index), or silica
microspheres (cf. Ref. [21, 23] and citations therein) where atoms close to the dielectric
interface interact evanescently with WGMs whose angular momentum according
to Eq. (31) is m ~ 10^{3}. At all levels of microcavity optics – quantum or classical,
linear or nonlinear – the fundamental task is to establish a modal basis, and the
methods we sketched in this chapter accomplish this in a way that provides physical
insight.

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