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Optics Express

  • Editor: C. Martijn de Sterke
  • Vol. 19, Iss. 12 — Jun. 6, 2011
  • pp: 11242–11255
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Nonlinear atom-field dynamics in high-Q cavities: from a BEC to a thermal gas

Tobias Grießer and Helmut Ritsch  »View Author Affiliations


Optics Express, Vol. 19, Issue 12, pp. 11242-11255 (2011)
http://dx.doi.org/10.1364/OE.19.011242


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Abstract

A cold gas of polarizable particles moving in the optical potential of a standing wave high finesse optical resonator acts as a dynamic refractive index. For a sufficiently strong cavity pump the optical forces generated by the intra cavity field perturb the particles phase space distribution, which shifts the optical resonance frequency and induces a nonlinear optical response. By help of the corresponding Vlasov equation we predict that beyond the known phenomenon of optical bi-stability one finds regions in parameter space, where no stable stationary solution exists. The atom field dynamics then exhibits oscillatory solutions converging to stable limit cycles of the system. The linearized analytical predictions agree well with corresponding numerical solutions of the full time dependent equations and first experimental observation in both cases.

© 2011 OSA

1. Introduction

Based on the surprisingly fast progress in the experimental technology of laser cooling and manipulation of dilute atomic gases as well as optical resonator and laser stabilization technology, it is now experimentally possible to confine larger and larger numbers of colder and colder atoms within the mode volume of stabilized very high finesse resonators [1

1. D. Kruse, M. Ruder, J. Benhelm, C. von Cube, C. Zimmermann, P. W. Courteille, T. Elsässer, B. Nagorny, and A. Hemmerich, “Cold atoms in a high-Q ring cavity,” Phys. Rev. A 67, 051802 (2003). [CrossRef]

,2

2. M. Khudaverdyan, W. Alt, I. Dotsenko, T. Kampschulte, K. Lenhard, A. Rauschenbeutel, S. Reick, K. Schörner, A. Widera, and D. Meschede, “Controlled insertion and retrieval of atoms coupled to a high-finesse optical resonator,” New J. Phys. 10, 073023 (2008). [CrossRef]

]. For slow enough particles the optical dipole force induced by the intracavity field significantly influences the atomic motion even in the dispersive limit at very large detunings, where absorption and spontaneous emission only play a minor role. In this limit the dispersive scattering of the particles simply acts as a dynamic refractive index changing the intracavity field evolution [3

3. P. Domokos and H. Ritsch, “Mechanical effects of light in optical resonators,” J. Opt. Soc. Am. B 20, 1098–1130 (2003). [CrossRef]

5

5. P. Domokos, P. Horak, and H. Ritsch, “Semiclassical theory of cavity-assisted atom cooling,” J. Phys. B 34, 187–198 (2001). [CrossRef]

]. The resulting complex coupled atom-field dynamics leads to a wealth of interesting physical phenomena and applications [6

6. J. M. Zhang, F. C. Cui, D. L. Zhou, and W. M. Liu, “Nonlinear dynamics of a cigar-shaped bose-einstein condensate in an optical cavity,” Phys. Rev. A 79, 033401 (2009). [CrossRef]

, 7

7. S. Ritter, F. Brennecke, K. Baumann, T. Donner, C. Guerlin, and T. Esslinger, “Dynamical coupling between a Bose–Einstein condensate and a cavity optical lattice,” Appl. Phys. B 95, 213–218 (2009). [CrossRef]

], which can be analysed theoretically using a wide range of models of very different complexity.

It has been known theoretically and experimentally for some time now, that the particles in this case act as a nonlinear optical medium through their motional response to the field, even though they are only weakly excited in the regime of linear polarizability at this frequency and intensity [11

11. S. Gupta, K. L. Moore, K. W. Murch, and D. M. Stamper-Kurn, “Cavity nonlinear optics at low photon numbers from collective atomic motion,” Phys. Rev. Lett. 99, 213601 (2007). [CrossRef]

, 12

12. A. Vukics, W. Niedenzu, and H. Ritsch, “Cavity nonlinear optics with few photons and ultracold quantum particles,” Phys. Rev. A 79, 013828 (2009). [CrossRef]

]. Even for the conceptually most simple case of a single mode standing wave resonator, the coupled atom-field system can exhibit optical bistability. For a BEC in such a cavity self pulsing solutions were experimentally found [13

13. F. Brennecke, S. Ritter, T. Donner, and T. Esslinger, “Cavity optomechanics with a Bose–Einstein condensate,” Science 322, 235–238 (2008). [CrossRef] [PubMed]

]. They can be qualitatively well described via a fairly simple two-mode expansion for the atomic mean field dynamics as a nonlinearly coupled oscillator model [14

14. D. Nagy, P. Domokos, A. Vukics, and H. Ritsch, “Nonlinear quantum dynamics of two BEC modes dispersively coupled by an optical cavity,” Eur. Phys. J. D 55, 659–668 (2009). [CrossRef]

]. This model was shown to be mathematically equivalent to a typical optomechanical setup of a high finesse resonator with a single field mode and a movable mirror at one end. Here we study the appearance of a very similar behavior for a general thermal gas in resonator and show that we can treat both limits of very low and high temperatures in a unified mean field model.

The paper is organized as follows. After presenting our general model in section 2 we use solve the Vlasov equation to find stationary states in section 3, which can give several solutions for a single choice of parameters. Using a linear perturbation analysis we then check the stability regions of these solution in section 4. In section 5 we construct and analyze the long time limit cycle solutions in the unstable regime. These are confirmed by numerical solutions of the time dependent Vlasov equation [Eq. (64)].

2. Model and basic equations

Fig. 1 Schematic drawing of the physical system. The particles are indicated as the shaded cloud in the center. The resonator mode decays at a rate κ and is driven by a laser characterised by the pump parameter η.

3. Steady states

Let’s turn to the determination of the steady states (ϕ 0,α 0) of Eqs. (5)(6). One easily finds
ϕ0(v)=vRU0|α0|241kvΔF0Δv,
(7)
where we have defined
ΔF0Δv:=F0(v+vR)F0(vvR)2vR.
(8)
Substituting Eq. (7) into Eq. (6) we get
0=(κ+iδ)α0ivRNU028k|α0|2α01vΔF0Δvdv+η,
(9)
where we have set δ := ΔcNU 0/2. Whenever the temperature is nonzero, we can define a function F(x), such that
F(x):=vTF0(xvT),
(10)
or equivalently F0(v)=vT1F(v/vT). To shorten the notation, let us introduce the factor
J(vR/vT):=121xF(x+vR/vT)F(xvR/vT)vR/vTdx,
(11)
and let us define the steady state photon number I 0 = |α 0|2. From Eq. (9) one gets a cubic equation for the steady state photon number
I0[κ2+(δI0ωRNU02J4ωT2)2]=η2.
(12)
Here, ωR := h̄k2/2m denotes the recoil frequency and ωT = kvT. Clearly, Eq. (12) can have up to three distinct real solutions. A typical response curve is depicted in Fig. (2). In the interval of effective detuning designated by the letter A the system allows for several possible steady states at once. Which one of these (if any) can be attained? A general investigation of the linear stability properties of steady states will shed some light on that question. Let us therefore establish the necessary and sufficient conditions for a given steady state solution (ϕ 0, I 0) to be stable or unstable.

Fig. 2 Normalized steady state photon number versus effective detuning δ/κ for N = 105, U 0 = 0.04κ, η =18κ, ωT = κ ωR = 0.5 · 10−3 κ and a Gaussian velocity distribution.

4. Stability analysis

4.1. Small Perturbation Analysis

The starting point of our analysis will be Eqs. (5)-(6), linearized around a steady state solution. Hence let us write
ϕ(v,t)=ϕ0(v)+δϕ(v,t),
(13)
α(t)=α0+δα(t).
(14)
Substituting into the equations of motion and keeping only terms that are linear in the deviations δϕ and δα, we obtain
δϕt+2ikvδϕ=ivRU02(α0δα*+α0*δα)ΔF0Δv,
(15)
δα˙=(κ+iΔ)δαiNU02α0Re{δϕ(v,t)}dv,
(16)
where we have set Δ:=δI0ωRNU02J4ωT2. It is advantageous to introduce the Laplace transforms with respect to time of δϕ, δϕ *, δα and δα *. To keep the notation to a minimum, we will not introduce new symbols for the transformed quantities. Introducing the “susceptibility”
χ(s):=NU02vR8(ΔF0/Δvs+2ikvΔF0/Δvs2ikv)dv,
(17)
we obtain the linear system of equations:
(s+κiΔI0χ(s)α02χ(s)α0*2χ(s)s+κ+iΔ+I0χ(s))·(δαδα*)=(I1I2),
(18)
where (I 1, I 2) denotes an initial value contribution. Let D(s) designate the determinant of the matrix on the left hand side of this last equation, i.e.,
D(s)=(κ+s)2+Δ22iΔI0χ(s).
(19)
We will henceforth call the complex valued function D(s) dispersion relation. As explained e.g. in [15

15. D. C. Montgomery, Theory of the unmagnetized plasma (Gordon & Breach, 1971).

], instability is equivalent to the existence of zeros of the dispersion relation with positive real part.

4.2. Limit of zero temperature-BEC

Before we direct our attention to the limiting case of a classical gas, let us now look closer at the opposite limit of a condensate at zero temperature. In that case, the velocity distribution is given by
F0(v)=δ(v),
(20)
where δ (v) is the dirac delta function. The steady-state photon number can be calculated from Eq. (12) with the replacements: J → − 1 and ωT → 2ωR. In order to make contact with the method of analysis used in other works, we decompose the solution to Eq. (5) according to
ϕ(v,t)=ϕ+(t)δ(vvR)+ϕ(t)δ(v+vR)+ϕ0(v,t),
(21)
where ϕ 0 is a solution of the homogeneous equation. This last contribution can be neglected for sufficiently smooth initial conditions due to rapid “phase mixing”, i.e.,
ϕ0(v,0)e2ikvtdv0,ast.
(22)
Defining x(t) := Re {ϕ + (t) +ϕ (t) } one easily checks that Eq. (5) leads to
x¨+(4ωR)2x=ωRU0|α|2,
(23)
α˙=[k+i(ΔcNU0/2)]αiNU02αx+η,
(24)
which are precisely the equations used to study the dynamics of a BEC inside a standing wave cavity using a Gross-Pitaevskii equation approach [16

16. F. Brennecke, T. Donner, S. Ritter, T. Bourdel, M. Köhl, and T. Esslinger, “Cavity QED with a Bose–Einstein condensate,” Nature 450, 268–271 (2007). [CrossRef] [PubMed]

]. The stability of steady states is governed by Eq. (19), in which the susceptibility takes on the form
χ(s)=iωRNU0221s2+(4ωR)2.
(25)
Without going into the detailed consequences of this susceptibility, we note that it can be demonstrated that no steady state with Δ=δ+NU02I032ωR>0 is stable. This also means that a positive effective detuning, δ > 0, implies instability.

4.3. Thermal gas - classical limit

Let us now investigate the dispersion relation for a nonzero temperature. Note that as it stands, Eq. (19) together with Eq. (38) are defined only for Re(s) > 0. To find the boundary that separates unstable from stable states, we need to find the limiting form of the dispersion relation as Re(s) → 0+. To this end we recall the Plemjel formula
limɛ0+1xy±iɛ=Pxyiπδ(xy).
(26)
Here and elsewhere, P denotes the cauchy principal value. Setting s = γ + and writing sloppily χ(ω) = limγ →0+ χ(s), we find
χ(ω)=NU02vR8k[1iPvΔF0/Δv(ω/2k2)v2dv+π2(ΔF0Δv|ω2kΔF0Δv|ω2k)].
(27)
It does not make much sense to try to extract general (i.e. true for arbitrary F 0) consequences of Eqs. (19) and (27), for even if we could analytically find the stability boundary in terms of a function I0crit=I0crit(δ,N,U 0,...), we would still have to find the intersection of this curve with the one specifying the actually possible combinations (I 0,δ,...) (i.e., Eq. (12)), and this is obviously a hopelessly complicated task. We will therefore concentrate on the physically most relevant velocity distribution, namely the thermal distribution. Well above the condensation threshold, it is given by the Maxwell-Boltzmann distribution given by
F0(v)=1vTπe(v/vT)2;J=2.
(28)
In this regime, we have vTvR, such that we may replace
ΔF0ΔvdF0dv
(29)
in all relevant equations, including Eq. (5). Finally, we should state the condition that the deviation of the particle distribution from the homogeneous be small, namely
ħU0|α|2mvT22,or4U0η2ωRκ2ωT21,
(30)
which simply means that the potential energy must be much smaller than the kinetic energy. We will not try to solve the stability problem fully analytically (which is impossible), but use graphical and numerical methods where necessary instead. Using Eq. (27) and writing limγ →0+ D(s) = Dr(ω) + iDi(ω), we find
Dr(ω)=κ2+Δ2ω2+NU02ΔI0ωRωT2(1ωωTDa(ω2ωT)),
(31)
Di(ω)=2κωπNU02ΔI0ωR2ωT2ωωTe(ω2ωT)2.
(32)
Here, Da(x) denotes Dawson’s integral which is defined as
Da(x)=ex20xet2dt.
(33)
Note that the dispersion relation belonging to a thermal velocity distribution can now be seen (upon substituting Δ → δ and U02I04η2) as being identical to the one considered in [8

8. T. Grießer, H. Ritsch, M. Hemmerling, and G. Robb, “A Vlasov approach to bunching and selfordering of particles in optical resonators,” Eur. Phys. J. D 58, 349–368 (2010). [CrossRef]

]. Therefore all the results exactly carry over to the present case: A steady state I 0 is unstable if and only if we have
Dr(ω0)>0
(34a)
for all ω 0 such that
Di(ω0)=0.
(34b)

4.4. Graphical solution of the stability problem

In order to extract the information included in the general criterion Eq. (34), we will resort to graphical means. To do so, we note that the two conditions Di = 0, Dr = 0 define a function
I0=I0crit(Δ)
(35)
(the dependence on the other parameters is understood implicitly), with the property that a steady state I 0 is unstable if and only if I0>I0crit. Hence (Δ, I0crit(Δ)) defines the stability boundary in the parameter-space spanned by (Δ,I 0). Using that
δ=ΔNU02ωR2ωT2I0,
(36)
we can parametrically plot the stability boundary, i.e. the mapping
Δ(ΔNU02ωR2ωT2I0crit(Δ),I0cirt(Δ))
(37)
in the space of interest, namely the one spanned by (δ,I 0). Together with the response curve defined by Eq. (12), this allows to analyze the stability properties of the system. Consider again the set of parameters used to plot Fig. (2). Figure (3) depicts the response curve and regions of instability defined by Eqs. (12) and (35) for two different values of the pump parameter. Clearly, there exists an interval of effective detunings, where two distinct steady photon numbers are stable (bistability). Perhaps more surprising however is the finding of an interval where there is no stable steady state at all (Fig. (3b))! To summarize the possible behavior of the system when varying the pump strength, we can state that below a certain value, the response curve is single-valued and lies entirely in the stable domain. Above that value, we find multi-valuedness and a region of bistability (Fig. (3a)). Only upon crossing a second threshold does the completely unstable interval appear (Fig. (3b)).

Fig. 3 Response curve (red) and regions of instability (shaded) for η = 13κ (a) and η = 18κ (b). Those parts of the response curve that lie inside the unstable region (red dashed), correspond to linearly unstable steady states. The black dashed line separates the parameter space into positive and negative Δ. The intervalls designated A correspond to bistability, the intervall designated B supports no stable steady state at all. All other parameters as in Fig. (2).

5. Long-time behavior

5.1. Classification

Whenever a steady state (two steady states) is (are) available, we expect the system to relax to it (or one of them) starting from arbitrary initial conditions. How does the system behave for parameters that allow for no stable steady state? To answer this question and to see whether the first guess is correct, we discreetized and solved Eqs. (5)-(6) numerically. Figure (4)a depicts the time averaged photon number (divided by the maximally possible) for different effective detunings for the usual set of parameters. It is clearly visible that this averaged photon number deviates from the steady state photon number only for detunings within the unstable interval, confirming our calculations. To our surprise and for generic sets of parameters, the system exhibits limit cycle oscillations within the completely unstable interval. Unlike in the case of a collisionless BEC, where oscillatory solutions depend on the initial conditions, these are true limit cycles independent of the initial conditions. This difference can be attributed to the phenomenon of “phase mixing” of the infinite number of oscillators in the classical case. The “observed” frequencies of the latter versus detuning are shown in figure (4)b. For the chosen parameters, these frequencies are of the order of the thermal frequency. For weak instability, the limit cycle oscillation is almost monochromatic. Figure (5) depicts these oscillations of the photon number for two detunings inside the unstable interval. Associated to these oscillations is a standing density wave formed by the particles.

Fig. 4 a: Response curve (stable parts red, unstable parts black dotted) and time averaged normalized photon number (blue circles) versus effective detuning. A deviation from the response curve occurs only inside the unstable intervall [−3.3,0], where the system exhibits limit cycle oscillations. b: Limit cycle frequency versus detuning. Parameters as usual.
Fig. 5 (a) Photon number as a function of time for δ = −κ (upper row) and δ = −3κ (lower row) and (b) normalized power spectrum. (c) Shape of the limit cycle. All additional parameters as in Fig. (2).

In Fig. (6) we compare the numerical solution of the reduced model with the solution of the Vlasov equation [Eq. (64)] for the same parameters as in Fig. (5). The agreement between the two models is excellent, confirming the validity of the approximations we introduced.

Fig. 6 Comparison between the reduced (blue) and the Vlasov model Eq. (64) (red). Apart from a phase shift, the results are almost identical. All parameters as in Fig. (2) and Fig. (5).

5.2. Limit cycles frequencies close to threshold

Fig. 7 Response curve and numerically determined limit cycle frequency (blue triangles) vs. theoretical frequency from Eq. (53) (green dashed). The blue circles show the time averaged photon number. Parameters: N = 105, kvT = κ, ωR = 10−3 κ and U 0 = 0.05κ.

6. Conclusions and outlook

The nonlinear dynamical response of a high-Q optical cavity filled with a cold gas of polarizable particles can be treated by an effective Vlasov type mean field approach over a large range of temperatures from close to zero (BEC) to a thermal gas. In both cases density waves of the medium are strongly coupled to the intra-cavity field dynamics. Besides the well known optical bi-stability effect concurrent with a hysteresis effect, we can clearly identify the regions in parameter space, where no stable stationary solutions exists and time evolution converges to a periodic limit cycle. Its frequency is characteristic for the temperature and gas properties and can be explicitly calculated in the BEC limit or close to threshold for a thermal gas. In both cases such oscillations have been reported experimentally and thoroughly studied [13

13. F. Brennecke, S. Ritter, T. Donner, and T. Esslinger, “Cavity optomechanics with a Bose–Einstein condensate,” Science 322, 235–238 (2008). [CrossRef] [PubMed]

, 17

17. M. Cristiani and J. Eschner (personal communication, 2011).

, 18

18. T. Valenzuela, M. Cristiani, H. Gothe, and J. Eschner, “Cold Ytterbium atoms in high-finesse optical cavities: cavity cooling and collective interactions,” in “Lasers and Electro-Optics 2009 and the European Quantum Electronics Conference. CLEO Europe-EQEC 2009. European Conference on,” (IEEE, 2009), p. 1. [CrossRef]

]. The observations agree well with our model. Interestingly, the effect is predicted to persist at much higher temperatures creating a nonlinear response in a linear medium. To observe it in practice at high temperatures would require sufficiently high intensities and densities or very high Q cavities. Here collisions can no longer be ignored, which can obscure this effect or even lead to a qualitative similar behavior based on a quite different physical mechanism. The collisionless model is in general a valid approximation for laser cooled thermal gases but gets more doubtful in the BEC limit. In the current experiments on this system at ETHZ [7

7. S. Ritter, F. Brennecke, K. Baumann, T. Donner, C. Guerlin, and T. Esslinger, “Dynamical coupling between a Bose–Einstein condensate and a cavity optical lattice,” Appl. Phys. B 95, 213–218 (2009). [CrossRef]

] or Tübingen their influence could largely be neglected so far. In the long run the system thus represents a well controllable and experimentally implementable toy system to study generic nonlinear dynamics in the transition range from classical to quantum mechanics. Adding extra modes or atomic species will allow to extend its scope and complexity in many directions.

7. Appendix: quantum and classical mean field limit

Here we show, how the mean field Wigner equation can be obtained from the quantum mechanical equations for N → ∞ polarizable particles coupled to one or more resonator modes. When spontaneous emission is neglected, the Heisenberg equation of motion for the particle field operator ψ^(x,t) reads
tψ^=iħ2m2x2ψ^+1iħV^ψ^,
(55)
wherein = (x, â 1,..., âM) = is the potential at x created by the photons of the M modes. Clearly we have [,ψ^] = 0. Let us introduce the phase space density operator
f^(x,p)=12πħeizp/ħψ^(xz2)ψ^(x+z2)dz,
(56)
where p = mv is the momentum. It may be noted that 〈 (x,p)〉 is the Wigner transform of the one body reduced density matrix ρp ,1(x,x′). It may be shown from Eq. (55) that the phase space density operator satisfies:
f^t+pmf^x+U^(x,q)f^(x,pħq)dq=0,
(57)
where
U^(x,q):=i2πħdz[V^(x+z/2)V(xz/2)]eiqz.
(58)
In case of a single mode of a standing wave resonator, we have (x,â) = h̄U 0 â â cos2(kx) and with this one readily finds from Eqs. (57) and (58) that the equation for the phase space density operator thus reads:
f^t+pmf^x+U02a^a^sin(2kx)[f^(x,p+ħk)f^(x,pħk)]=0.
(59)
This equation is to be supplemented by the Heisenberg-Langevin equation for the annihilation operator â of the mode:
a^˙=(κ+iδ)a^iU0a^0Ldxf^(x,p)cos(2kx)dp+η+a^in.
(60)
The mean field limit N ≫ 1 can be found by decomposing âj = αj + δâj and = f + δf̂ with 〈δâj〉 = 〈δf̂〉 = 0, taking the average of Eq. (57) and neglecting all correlations. This procedure leads to Wigner’s equation (sometimes called the Quantum Vlasov equation) for the expectation value of (x, p)
ft+pmfx+U02|α|2sin(2kx)[f(x,p+ħk)f(x,pħk)]=0,
(61)
α˙=(κ+iδ)αiU0α20Ldxf(x,p)cos(2kx)dp+η.
(62)
If we further assume that the typical momentum is much larger than h̄k, we can develop
f(x,p±ħk)f(x,p)±ħkfp(x,p)+ħ2k222fp2(x,p).
(63)
Then Eq. (61) becomes
ft+pmfx+ħkU0|α|2sin(2kx)fp=0,
(64)
which is Vlasov’s equation. Hence, the classical model employed in [8

8. T. Grießer, H. Ritsch, M. Hemmerling, and G. Robb, “A Vlasov approach to bunching and selfordering of particles in optical resonators,” Eur. Phys. J. D 58, 349–368 (2010). [CrossRef]

] and in this work is expected to be accurate if kBT 2 k 2/2m, i.e. if the thermal energy (or kinetic energy per particle) is much higher than the recoil energy.

Acknowledgments

This work was supported by the Austrian Science Fund FWF under grant Nr. I119-N16 and SFB F40 FOQUS. We thank Claus Zimmerman for stimulating discussions and Matteo Cristiani and Jürgen Eschner for helpful comments and early communication of their experimental results.

References and links

1.

D. Kruse, M. Ruder, J. Benhelm, C. von Cube, C. Zimmermann, P. W. Courteille, T. Elsässer, B. Nagorny, and A. Hemmerich, “Cold atoms in a high-Q ring cavity,” Phys. Rev. A 67, 051802 (2003). [CrossRef]

2.

M. Khudaverdyan, W. Alt, I. Dotsenko, T. Kampschulte, K. Lenhard, A. Rauschenbeutel, S. Reick, K. Schörner, A. Widera, and D. Meschede, “Controlled insertion and retrieval of atoms coupled to a high-finesse optical resonator,” New J. Phys. 10, 073023 (2008). [CrossRef]

3.

P. Domokos and H. Ritsch, “Mechanical effects of light in optical resonators,” J. Opt. Soc. Am. B 20, 1098–1130 (2003). [CrossRef]

4.

P. Horak, G. Hechenblaikner, K. M. Gheri, H. Stecher, and H. Ritsch, “Cavity-induced atom cooling in the strong coupling regime,” Phys. Rev. Lett. 79, 4974–4977 (1997). [CrossRef]

5.

P. Domokos, P. Horak, and H. Ritsch, “Semiclassical theory of cavity-assisted atom cooling,” J. Phys. B 34, 187–198 (2001). [CrossRef]

6.

J. M. Zhang, F. C. Cui, D. L. Zhou, and W. M. Liu, “Nonlinear dynamics of a cigar-shaped bose-einstein condensate in an optical cavity,” Phys. Rev. A 79, 033401 (2009). [CrossRef]

7.

S. Ritter, F. Brennecke, K. Baumann, T. Donner, C. Guerlin, and T. Esslinger, “Dynamical coupling between a Bose–Einstein condensate and a cavity optical lattice,” Appl. Phys. B 95, 213–218 (2009). [CrossRef]

8.

T. Grießer, H. Ritsch, M. Hemmerling, and G. Robb, “A Vlasov approach to bunching and selfordering of particles in optical resonators,” Eur. Phys. J. D 58, 349–368 (2010). [CrossRef]

9.

J. Javaloyes, M. Perrin, G. L. Lippi, and A. Politi, “Self-generated cooperative light emission induced by atomic recoil,” Phys. Rev. A 70, 023405 (2004). [CrossRef]

10.

R. Bach, K. Burnett, M. d’Arcy, and S. Gardiner, “Quantum-mechanical cumulant dynamics near stable periodic orbits in phase space: application to the classical-like dynamics of quantum accelerator modes,” Phys. Rev. A 71, 33417 (2005). [CrossRef]

11.

S. Gupta, K. L. Moore, K. W. Murch, and D. M. Stamper-Kurn, “Cavity nonlinear optics at low photon numbers from collective atomic motion,” Phys. Rev. Lett. 99, 213601 (2007). [CrossRef]

12.

A. Vukics, W. Niedenzu, and H. Ritsch, “Cavity nonlinear optics with few photons and ultracold quantum particles,” Phys. Rev. A 79, 013828 (2009). [CrossRef]

13.

F. Brennecke, S. Ritter, T. Donner, and T. Esslinger, “Cavity optomechanics with a Bose–Einstein condensate,” Science 322, 235–238 (2008). [CrossRef] [PubMed]

14.

D. Nagy, P. Domokos, A. Vukics, and H. Ritsch, “Nonlinear quantum dynamics of two BEC modes dispersively coupled by an optical cavity,” Eur. Phys. J. D 55, 659–668 (2009). [CrossRef]

15.

D. C. Montgomery, Theory of the unmagnetized plasma (Gordon & Breach, 1971).

16.

F. Brennecke, T. Donner, S. Ritter, T. Bourdel, M. Köhl, and T. Esslinger, “Cavity QED with a Bose–Einstein condensate,” Nature 450, 268–271 (2007). [CrossRef] [PubMed]

17.

M. Cristiani and J. Eschner (personal communication, 2011).

18.

T. Valenzuela, M. Cristiani, H. Gothe, and J. Eschner, “Cold Ytterbium atoms in high-finesse optical cavities: cavity cooling and collective interactions,” in “Lasers and Electro-Optics 2009 and the European Quantum Electronics Conference. CLEO Europe-EQEC 2009. European Conference on,” (IEEE, 2009), p. 1. [CrossRef]

OCIS Codes
(190.1450) Nonlinear optics : Bistability
(270.3100) Quantum optics : Instabilities and chaos

ToC Category:
Quantum Optics

History
Original Manuscript: January 19, 2011
Revised Manuscript: March 14, 2011
Manuscript Accepted: March 14, 2011
Published: May 25, 2011

Citation
Tobias Grießer and Helmut Ritsch, "Nonlinear atom-field dynamics in high-Q cavities: from a BEC to a thermal gas," Opt. Express 19, 11242-11255 (2011)
http://www.opticsinfobase.org/oe/abstract.cfm?URI=oe-19-12-11242


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References

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