Abstract

We propose a scheme for generation of a nonlinear coherent state (NCS) of a mechanical resonator (MR) in an optomechanical micro-cavity, in which a two-level quantum dot (QD) and the microcavity are respectively driven by a strong laser and a weak laser. This microcavity can be engineered within a photonic band-gap (PBG) material. By properly tuning the frequency of the weak driving field, two-photon blockade phenomenon occurs. The QD-cavity subsystem can evolve into a dark state due to the damping of the microcavity and the elimination of the decay rate of the QD at selected frequencies in the PBG material. In this situation, the phonon mode of the MR can be prepared into a NCS, which is a non-classical state and possesses the sub-Poisson statistics. We also demonstrate the Wigner function of the NCS, which negativity implies its non-classicality.

© 2016 Optical Society of America

1. Introduction

In the emerging field of optomechanical systems, numerous rapid progresses with the radiation pressure have recently been obtained during measuring and cooling the mechanical motion [1,2]. It is reported that the quantum control of mechanical motion and the control of light or microwave field in cavity optomechanical systems can be achieved to some extent [3, 4]. The cooling of a mechanical resonator (MR) to its ground state of motional degree of freedom is an important procedure [5, 6]. Experimentally, it has been realized that micro- and nanomechanical resonator can be cooled to or near its quantum ground state in various mechanical systems either through direct cryogenic cooling [7], or laser cooling using microwave [8, 9] and optical cavity fields [10]. When the MR is cooled to the ground state, it is possible to generate and detect the non-classical states of the MR, such as the squeezed state [3, 11, 12] and the entangled state [13]. The squeezed state of a movable mirror can be obtained via feeding broadband squeezed vacuum light into the cavity in dissipative optomechanics [11]. By driving an optomechanical cavity with two controllable lasers with different amplitudes, the dissipative mechanism with the driven cavity acts as an engineered reservoir, the squeezed state of a MR can also be obtained [14]. Recently, a scheme for generation of the non-classical state of the MR is proposed, in which the MR is nonlinearly coupled to a two-level system [15]. The non-classical states of the MR can also be produced by applying the generation of motional dark states [16].

The nonlinear coherent state (NCS) as another concerned non-classical states has also attracted many interests in view of its promising application in quantum algebras [17–19], which is based on noncommutative space rather than the usual space [20]. In the description of the motion of a trapped ion, Vogel et al. put forward the definition of the NCS as the eigenstate of the non-Hermitian operator f()a [21,22], where the deformation function f() is an operator-valued function of number operator = aa. In this case, the commutator [f(n)a, a f(n)] is not a constant or a linear function of the generators of the Lie algebra but nonlinear in the generators. It is found that the NCS exhibits non-classical features such as amplitude squeezing and self-splitting, which are accompanied by pronounced quantum interference effects [21,23]. The generation schemes of the NCS are explored in some other literatures, e.g. a single-atom laser [17], a micromaser under the intensity-dependent Jaynes-Cummings model [24], an exciton in a wide quantum dot [25], and a trapped ion [21,26]. The stationary state of a single-atom laser is shown to be a phase-averaged NCS and the interaction-induced transition probabilities of the single-atom laser effectively depend on the field intensity [17]. The quantized motion of an exciton in a wide quantum dot interacting with two laser beams can be prepared in a NCS and the exciton states can be utilized for quantum information processes [25]. The NCS is employed to discuss the optimization of quantum information which is important in binary (or multibinary) communication [27]. Usually, the discussion of NCS is related with the non-Gaussian quantum states. The occurrence of the negativity of the Wigner function (WF) for some non-Gaussian states implies that they possess the non-classical properties. The non-Gaussian state plays a significant role in entanglement distillation [28, 29], quantum computation with cluster states [30], loophole-free Bell tests with the continuous variables [31, 32], and quantum error correction [33]. Besides, the non-Gaussian state can be exploited to improve the quantum communication protocols in quantum information processing [34] and improve the optimal estimation of losses at the ultimate quantum limit [35]. Recent years, it is found that the optomechanical systems possess the intrinsic nonlinearity thus hold promising applications in quantum information processing [36] and ultra-precision measurement [37]. So how to prepare the NCS of a MR in the optomechanical system is a topic of great significance.

The generation of a non-classical light source is one of the major requirements of quantum information processor associated with modern semiconductor fabrication technology [38]. For this purpose, employing photon blockade phenomenon is one of methods. This phenomenon is first proposed by Imamoḡlu [39], which occurs when the absorption of a first input photon by an optical device blocks the transmission of a second one, thereby leading to non-classical output photon statistics. The inherent non-classical property at single photon level provides a way to control single photon, which is commonly used in quantum key distribution system [40] and plays a fundamental role in proposed devices for quantum information processing [38]. The photon blockade is first observed in cavity quantum electrodynamics(QED) systems with the strong atom-cavity coupling [41], and then has been demonstrated in different solid-state schemes, e.g. in cavity QED systems with a quantum dot (QD) in photonic crystal cavities [42], with a QD strongly coupled to a photonic crystal resonator [43] and other circuit-QED systems [44, 45]. Recently, much attention has been focused on multi-photon blockade [46–48]. The photon pairs generated by the blockade are often required for generation of the entangled light. Two-photon blockade in the strong-coupling qubit-cavity system is investigated using a modified Lindblad master equation [47]. Two- and three- photon blockade phenomena observed respectively in the standard cavity QED and circuit QED systems are attributed to photon-induced tunneling [46]. It is demonstrated that multi-photon blockade may inhibit additional photon being absorbed, which is caused by the nonlinearity of the Jaynes-Cummings (JC) model [48]. In a word, it makes sense to propose a scheme in the optomechanical system for preparing a non-classical state of the MR via the multi-photon blockade.

In this paper, we propose a scheme for generating the NCS of the phonons via the two-photon blockade and the dark state of the system, then investigate its non-classical properties. A two-level QD is driven by a strong laser and the microcavity is engineered within a photonic band-gap (PBG) material. In the dressed state basis induced by the strong field, resonant excitation of two-photon can be implemented by the weak driving field. In this case, the two-photon blockade can be realized in the optomechanical system by tuning the frequency of the weak driving field. The optomechanical system can evolve into a dark state due to the damping of the microcavity and the decay of the dressed QD at selected frequencies in the PBG material. As a result, the phonon mode of the MR can be prepared in a NCS, which is a non-classical state and possesses the properties of the sub-Poisson statistics. We also demonstrate the Wigner functions (WF) of NCS, which negativity implies its non-classicality.

The paper is organized as follows. In section 2, we show the physical model for the driven QD inside an optomechanical microcavity coupled to a weak laser. The whole system is transformed into the polaron frame, and the master equation of the optomechanical system is obtained using the rotating wave, secular and Born-Markov approximations. The preparation of the NCS and its properties are investigated in section 3. Finally, we draw our conclusions in section 4.

2. System and master equations

2.1. Description of the system

The system under consideration consists of a high-Q optomechanical microcavity, which is technically engineered inside a PBG material and coupled with an excitonic two-level QD. The optomechanical microcavity with the frequency ωc can be coupled with the MR of frequency ωm via a combination of radiation pressure and photostriction [49]. The cavity is also driven by a weak laser field with frequency ωp and amplitude αp, as shown in Fig. 1. The two-level QD is represented by the usual Pauli spin- 12 operators σ+ and σ, which satisfy the commutation relations [σ+, σ] = σz and [σz, σ±] = ±2σ±. It is placed inside the microcavity and coupled to another strong laser field with frequency ωL and Rabi frequency Ω. This model can be realized by a optomechanical microcaviy, which includes a defective photonic crystal nanobeam and a MR with the “breathing” mode [50], and a QD ironing the surface of the beam (see Fig. 1). The motion of the MR manifests as a successive of contractions and expansions along the x axis and is driven by a strong laser. The microcavity is driven another weak laser light. In the rotating frame at the frequency of ωL, the total Hamiltonian of the system is given by (setting = 1 throughout the paper)

H=H0+HI,
in which the term
H0=Δa2σz+Ω(σ++σ)+Δcaa+ωmbb+λΔλaλaλ,
describes the unperturbed Hamiltonian of the coherently driven QD under the rotating-wave approximation, the microcavity, the phonon and the photonic crystal vacuum reservoir. The operators a(a), b(b) and aλ(aλ) are the annihilation (creation) operators of the microcavity field, the mechanical mode and the photonic crystal vacuum reservoir, respectively. Δj = ωjωL(j = a, c, λ) are the detunings of the QD’s resonance frequency, of the microcavity-mode frequency and of the crystal reservoir frequency from the driving field frequency, respectively. Under the electric dipole and the rotating-wave approximations, the interactions between the driven QD, the microcavity, the mechanical mode and independent bathes can be expressed as [51]
HI=[αpaeiΔpt+g1aσ+λgλ(ωλ)aλσ+h.c.]+g2aa(b+b),
where Δp = ωpωL is the frequency difference between the weak (ωp) and strong (ωL) driving field. The coefficients g1 and g2 describe the coupling strengths of the microcavity with the QD and phonon mode of the MR. We have assumed that the coupling coefficient g1 is independent on frequency. g2 = ωcx0/L in Eq. (3) is the single-photon coupling strength, where x0=12Mωm is the zero-point fluctuation of the MR with its mass M, and microcavitys rest length is L.

 

Fig. 1 Schematic of the considered system. A photonic crystal nanobeam is clamped at both ends. The optomechanical microcavity with the frequency ωc can be coupled with the MR of frequency ωm via a combination of radiation pressure and photostriction [49]. The MR can travel fast and alternating expansions and contractions along the x axis, which we refer to as a “breathing” mode [50]. A QD (the small red dot in Fig) driven by a strong laser is within the photonic crystal microcavity.

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In Eq. (3), the information about the frequency dependent of the photonic crystal is stored in the coupling strength gλ (ωλ) which can be written as gλ (ωλ) = gλD(ωλ) with gλ being a constant proportional to the dipole moment of the QD and D(ωλ) corresponding to the transfer function of the photonic crystal reservoir. Experimentally, when the length of the waveguide in a 3D heterostructure is 12μm, the local photon density of states (LDOS) is demonstrated to have a jump of an order of 100 within a range of relative frequencies about 10−4 ωb, where ωb is the cutoff frequency of the waveguide mode [52]. The large discontinuity with step-shaped LDOS can be provided by a waveguide cutoff mode within a 3D-2D PGB heterostructure [52,53]. Besides, a forklike shape LDOS can also be engineered using a trimodal waveguide architecture in a PBG microchip using suitable cutoff frequencies for two of the three waveguide modes [52]. In this paper, the large discontinuity in the LDOS and the relative positions of the relevant frequencies considered in our study are presented in Fig 2, such that |D(ωλ)|2 = u(ωλωb), u(ωλωb) being a unit step function. One valuable feature of waveguide structure is that near the waveguide cutoff frequency, the field is almost the same for rods along the waveguide direction. Therefore, the QD which is close to rods and along the waveguide direction will experience the same field. The cavity resonant frequency can be engineered to be within a PBG or high-LDOS region and near the LDOS discontinuity [52, 54]. Since the QD is driven by a strong laser field, three Mollow bands appear in the fluorescent scattering spectrum [55]. When the Mollow components at different frequencies (especially at lower and higher frequencies) fall within the high LDOS region or the low LDOS region, then they will experience strongly different mode densities [53]. It will strongly influence the dynamics of QD and results in new emission effect, i.e. |D(ωλ)|2 = 1 for ωλ > ωb and |D(ωλ)|2 = 0 for ωλ < ωb. This technique supplies a way for us to eliminate spontaneous transitions at selected frequencies of the QD-cavity system. In our scheme, this property of the PBG material is very important to generate the NCS of the MR, because the generation of the dark state of the system requires the elimination of the spontaneous transitions at both lower and central frequencies of the QD-cavity system, as will be discussed later.

 

Fig. 2 Schematic representation of the LDOS and the relative position of the relevant frequencies considered in our paper. ωb, ωc andωL are the photonic density-of-states bandedge frequency, the cavity frequency and the coherent driving frequency respectively. The vertical arrows represent the emission lines of the Mollow spectrum of the QD resonance fluorescence which central band occurs at the frequency ωL, and the side bands components at ωL ± 2Ω̄.

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The strong driving field can be viewed as a dressing field for the QD. Therefore, we diagonalize the first two terms of the Hamiltonian H0, i.e., the Hamiltonian of the QD and the coherent interaction with the strong driving field, to find the eigenstates of the coherently driven QD [56]

|+=s|g+c|e,|=c|gs|e,
where the parameters c=12+Δa4Ω, s=1c2. Ω¯=Ω2+(Δa2)2 is the effective Rabi frequency. In this dressed-state basis induced by the strong driving field, the unperturbed Hamiltonian H0 and the interaction Hamiltonian HI in Eq. (1) are rewritten as
H0=Ω¯R3+Δcaa+ωmbb+λΔλaλaλ.
HI={[g1a+λgλ(ωλ)aλ](scR3+c2R+s2R+)+h.c.}+αp(aeiΔpt+aeiΔpt)+g2aa(b+b),
where Rij = |i〉 〈j|(i, j = +, −) are the dressed-state transition operators, R3 = R++R−− is the population difference operator.

In order to conveniently study the dynamics of the system including QD, microcavity and the MR, we transform the total system into the polaron frame [57] and introduce a renormalized MR-microcavity coupling strength. The polaron transformation of the total system can be written as

H=eS(H0+HI)eS,
where S = ηaa(bb), and η = g2/ωm. The Lamb-Dicke parameter η measures the localization of the motional ground state of the MR relative to the (effective) wavelength of the transition. After the transformation, the Hamiltonian of the whole system can be written as
H=H0+HI1+HI2+HI3
with
H0=Ω¯R3+Δcaa+ωmbb+λΔλaλaλ,
HI1=g1aeη(bb)(scR3+c2R+s2R+),
HI2=αpaeη(bb)eiΔpt+h.c.,
HI3=λgλ(ωλ)aλ(scR3+c2R+s2R+)+h.c.
describing the Hamiltonian of a modified system (the dressed QD plus the optomechanical microcavity). The Hamiltonian H′I1 in Eq. (10) represents the interaction among the QD, the microcavity and the MR, H′I2 characterizes the interaciton of the microcavity with the MR and the weak laser field. It is indicated from Eq. (11) that the absorption processes of a cavity photon are inseparable from the transition processes of the QD and the generation or annihilation of phonons. In the following section, this will be employed to generate the NCS of the MR. The term H′I3 is the interaction of the QD with the photonic crystal reservoir.

2.2. Derivation of the effective master equation

By use of the Baker-Campbell-Hausdorff Relation [58, 59], the exponential eη(bb) can be expanded as

eη(bb)=eη22m,n=0ηm(η)nm!n!bmbn.
In this paper, we only focus on the case that the QD is red detuned from the microcavity, i.e., Δc = 2Ω̄ and ωm ≈ 4Ω̄. With this condition, in the interaction picture defined by H′0, we employ the rotating wave approximation to neglect the terms which oscillate rapidly and keep the nearly resonant oscillating terms. The term H′I1 in Eq. (10) is given by
H˜I1(t)=g1[c2aR+ηs2af1(bb)bR+eiδmt+h.c.],
where g1=g1eη22 and δm = Δc + 2Ω̄ − ωm is the detuning of the MR with the microcavity and the Rabi frequency of the strong driving field, which in our concerned case satisfies the condition δm ≪ Ω̄, ωm. The operator-valued function f1(bb) in Eq (14) is of the form
f1(bb)=n=0(1)n+1η2nn!(n+1)!bnbn.
It is obvious from Eq. (14) that a multi-particle system consisting of the QD, the microcavity and the mechanical mode remains to be investigated. The term H′I2 in Eq. (11) is of form
H˜I2(t)=αp[af2(bb)eiδcpt+h.c.],
in which αp=αpeη22 and δcp = Δc − Δp is the detuning of the microcavity from the weak laser field. In Eq. (16), we have used the condition δcpωm to employ the rotating wave approximation. This condition will be discuss later. Another operator-valued function f2(bb) in Eq. (16) is
f2(bb)=n=0(1)nη2nn!n!bnbn.
The term H′I3 is of the form
H˜I3=λgλ(ωλ)aλ[scR3eiΔλt+c2R+ei(Δλ2Ω¯t)s2R+ei(Δλ+2Ω¯)t]+h.c.

As mentioned above, in order to explore the dynamics of the system with a three-particle direct interaction, we find that there is one time-independent Hamiltonian in I1(t), i.e., HJ = g′1c2(aR−+ + aR+−), which corresponds to the Jaynes-Cummings(JC) interaction Hamiltonian describing the interaction between the dressed QD and the microcavity. The JC model has been known to describe an extensive range of coherent phenomena including the spontaneous collapse, the periodic Rabi-flopping and revivals of a two-level atom in a resonant cavity [60]. The non-classical character is investigated by the transmitted light in a driven cavity-atom system both in theoretical [61] and experimental [41] literatures. It is found that remarkable nonlinearities occurs when the coupling between the cavity mode and the atom is strong. HJ can be diagonalized, and the eigenstates consist of a ground state |0〉 = |−, 0〉 with energy E0 = 0 and excite states |E±(n)=12(|,n±|+,n1) (n = 1, 2,···) with corresponding energies E±(n)=±ng1c2. |+(−), n〉 denotes the higher- (lower-) energy eigenstate with n excitations. The splitting between the energy eigenstates in each manifold has a nonlinear dependence on n. We assume that the coupling coefficient g′1c2 is much larger than the damping rate κ of the microcavity. By cooperatively modulating the frequencies ωp, ωL of the driven fields and the strength Ω, two-photon transition processes can be tuned to occur resonantly and the other excitations are far off resonance. In this case, one can truncate the Hilbert space and describe the QD-cavity subsystem with the following four states

|0=|,0,|1=|,1|+,02,|2=|,1+|+,02,|3=|,2|+,12.

For the convenience of expression, in Eq. (19) we have denoted |0〉, |E(1), |E+(1) and |E(2) as |0〉, |1〉, |2〉 and |3〉 respectively, with corresponding energies E0 = 0, E1 = −g′1c2, E2 = g′1c2 and E3=2g1c2. In our scheme, the decay rate κ and γ of the coupled QD-cavity system are much larger than that of the MR. We declare here that the four states concerned in Eq. (19) are established in the subspace of the reduced QD-cavity subsystem, which involves the strong driving field ωL. Moreover, the resonant frequency of the cavity ωc is at least a factor of 103 with the frequency of ωm [8–10,50,62–66]. In this situation, the energy interval between the subspaces of the reduced QD-cavity system is much larger than the energy interval of the MR Fock space. Therefore, we suppose that these above four states are enough to describe the considered system even when taking 15 Fock states of the MR into account as shall be discussed in section 3. The energy diagram of the above states of QD-cavity subsystem is shown in Fig 3. As a result, the time-dependent Hamiltonian HJ are diagonalized as

HJ=i=03Ei|ii|.

 

Fig. 3 Schematic of the four-level model. The two-photon transition from the ground state |0〉 to the higher excited state |3〉 is resonant.

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In this new basis of Eq. (19), we can rewrite the photon creation operator of the cavity and the transition operator of the QD as

a=12(|1+|2)0|+|3(2+121|+2122|),R+=12|0(1|2|)12(|1+|2)3|.
Using Eq. (21) and its conjugate counterpart, one can rewrite the rest part of I1, I2 and I3. After that, we work in the interaction picture defined by HJ of Eq. (20) and obtain the interaction Hamiltonian of the system in the truncated Hilbert space as
VI=V1+V2,
where
V1=Ωe[|30|f1(bb)b+h.c.],
with the parameter Ωe=22ηs2g1. In the calculation of Eq. (23), we have set δm=2g1c2. The second term V2 in Eq. (22) is of the form
V2=αp[22|10|eiΔ1t+22|20|eiΔ2t+2+12|31|eiΔ3t+212|32|eiΔ4t]f2(bb)+h.c.
with the parameters being Δ1 = δcpg′1c2, Δ2 = δcp + g′1c2, Δ3=δcp(21)g1c2, and Δ4=δcp(2+1)g1c2. V2 represents the interaction with larger detuning between the microcavity and the MR when we assume the detuning as δcp=22g1c2αp. It is obvious that δm and δcp are the same order of magnitude. As stated before, since we are only interested in the case Δc = 2Ω̄ and ωm ≈ 4Ω̄, the parameter δm = Δc + 2Ω̄ − ωm satisfies the condition δm ≪ Ω̄, ωm. Consequently, the condition δcpωm is satisfied automatically, which is the condition we have used during the calculation of Eq. (16). The two states |1〉 and |2〉 in the Eq. (24) are intermediate states for the two-photon transition and they are largely detuned from the weak probe field. Therefore, we choose δcp=22g1c2 to adiabatically eliminate |1〉 and |2〉 and calculate the effective Hamiltonian of V2 as
V2=iV2(t)V2(t)dt.
In the derivation of the effective Hamiltonian, we also neglect some constant terms which correspond to the energy shift of the system. As a result, the effective Hamiltonian V′2 in the interaction picture takes the form
V2=Ωe[χf22(bb)|30|+h.c.]
with χ=4αp2ηg12s2c2. It is clearly that two-photon resonance can be induced by the weak driven field, because the QD-cavity subsystem can be excited from the ground state |0〉 to state |3〉 through the absorption of two-photon when two photons with frequency ωp from the weak field are present, as shown in Fig 3. It also should be mentioned that once this subsystem absorbs two photons with frequency ωp, the energy level is shifted far-off resonance with two-photon resonance frequency. It prevents another two photons from entering the microcavity until the previous two photons leave, which is known as two-photon blockade.

Taking the similar processes, we now calculate the dissipative terms. For the dissipative contribution originating from the interaction of the QD with the photonic crystal reservoir as depicted by I3 in Eq. (18), we assume that the coupling between the QD and photonic crystal vacuum reservoir is weak and its change on the reservoir is negligible. Also, a fast time scale for the decay of the reservoir correlation is assumed such that the secular [67] and Born-Markov [58,59] approximations can be used. In quantum optics, these standard approximations have been applied with considerable success to a great deal of atom-field interaction problems in photonic crystals [55, 68]. A similar method with details has been introduced in other literature [69]. We apply the second-order perturbation theory to trace over the reservoir degrees of freedom from the following equation

ρt=0tdtTrR{[H˜int(t),[H˜int(t),ρT(t)]]},
where TrR denotes tracing over the photonic crystal vacuum reservoir variables, and int (t) is the interaction of the QD with photonic crystal in the interaction picture, i.e., int (t) = I3(t). Under the Born-Markov approximation, the operator ρT (t′) in Eq. (27) can be replaced by ρ(t) ⊗ ρλ(0), where ρ(t) is the reduced density operator for the three-particle (QD, cavity and MR) system and ρλ(0) is the initial density operator of the photonic crystal vacuum reservoir. After straightforward calculations, we obtain the dissipative term λρ induced by the interaction of the QD with photonic crystal vacuum reservoir, which has Liouvillian form as
λρ=γ0𝒟[R3]+γ𝒟[R+]+γ+𝒟[R+],
The terms proportional to γ0,± in Eq. (28) are responsible for the damping among the dressed states of the dressed-QD system. The parameter γ0=2πc2s2λgλ2|D(ωλ)|2δ(ωλωL) is the spontaneous emission rate with which the electron of the QD transits at central frequency of the dressed states. The parameter γ+=2πc4λgλ2|D(ωλ)|2δ(ωλωL2Ω¯) represents spontaneous emission rate of electron which occurs at high frequency ω+ = ωL + 2Ω̄ of the dressed states from the upper dressed state |+〉 of one manifold to the lower dressed state |−〉 of the manifold below, whereas the parameter γ=2πs4λgλ2|D(ωλ)|2δ(ωλωL+2Ω¯) describes spontaneous emission which occurs at lower frequency ω = ωL − 2Ω̄ of the dressed states from the lower dressed state of one manifold to the upper dressed state of the manifold below. However, in our paper, the microcavity is engineered in a PBG material. As mentioned before, by appropriately engineering the system topology, the QD coupled to the cavity field can experiences the discontinuity of the LDOS [54]. It strongly influences the dynamics of QD and results in new emission effects, i.e. |D(ωλ)|2 = u(ωλωb) = 0 for ωλ < ωb and |D(ωλ)|2 = 1 for ωλ > ωb. For simplicity, we also assume that the photonic density of modes is constant over the spectral regions surrounding the dressed-state resonant frequencies ωL and ωL ± 2Ω. In this way, the spontaneous transitions of the QD at frequencies ω and ωL can be selectively eliminated, i.e., γ0 = γ = 0.

The decay of the cavity cρ = κ𝒟[a] is phenomenologically added with the definition 𝒟[𝒪]=12(2𝒪ρ𝒪ρ𝒪𝒪𝒪𝒪ρ). κ is the decay rate of the microcavity which may lead to a trend towards vacuum mode of the cavity. In fact, cρ = κ𝒟[a] can also be obtained following above procedures of λρ. After the polaron transformation of Eq. (7), the decay of the cavity is of the form

cρ=κ𝒟[aeη(bb)].
Expand the dissipative terms cρ and λρ into the truncated subspace of the dressed QD-cavity system as illustrated in Eq. (19), we finally obtain the master equation for the density operator ρ of the three-particle system as the form
ddtρ=i[V1+V2,ρ]+λρ+cρ,
with the dissipation terms λρ and cρ in Eq. (30) being the form
λρ=j=12{γ2𝒟[|j0|]+γ+2𝒟[|0j|]+γ4𝒟[|3j|]+γ+4𝒟[|j3|]+γ0{𝒟[|00|]+𝒟[|12|]+𝒟[|21|]},cρ=κeη24{j=12m,n,k=0ϒbmbn|0j|ρ|j0|bkbm+kn+j=12m,n,k=0Bjϒbmbn|j3|ρ|3j|bkbm+kn}κeη24j=12(1+Bj)(|jj|ρ+ρ|jj|),
where ϒ = ηm+l(−η)n+k/(m!n!k!l!), Bj=322(1)j. In the derivations of cρ and λρ, the intermediate states |1〉 and |2〉 (see figure 3) are not abandoned because they are sometimes occupied during the dissipation processes.

The purpose of this paper is to prepare the NCS of the MR’s phonon mode. In the next section, we will demonstrate the generation of the NCS applying the dark state of the total system. The properties of this non-classical state is also explored.

3. Preparation of NCS using dark state

Up to now, we have obtained the effective master equation which is able to describe the dynamics of three-particle system including the MR, the microcavity and the dressed QD. We also mentioned the occurrence of two-photon blockade. However, another significant object of this paper is to prepare the NCS of the phonon. In the following, on the basis of the effective Hamiltonian, we investigate the dark state of the system and propose a scheme to generate the NCS of the MR. Some properties of the NCS are also discussed.

We first consider the case that the interaction Hamiltonian V1 and V′2 can be neglected, i.e., V1 = V′2 = 0. In this case, the dynamic of the whole system is determined by the dissipation term ℒρ = λρ + cρ. The steady-state behavior of the system can be obtained via the steady equation ℒρ(∞) = 0 with ρ()=j=03ρjj|jj|ρbj(), where ρbj(∞) is the reduced density operator of the MR in the steady state when the QD-cavity subsystem is in state |j〉. ρjj = 〈j|ρ| j〉 are elements of the reduced density matrix of the QD-cavity subsystem. In the steady-state solution, we find that ρ00 = [γ+/(γ+ + γ)]2. It should be stressed here that if the spontaneous emission rate can be eliminated γ = 0, the reduced QD-cavity subsystem evolves into the state |0〉 = |−, 0〉 as depicted in Eq. (19), i.e., ρ00 = 1, ρjj = 0, (j = 1, 2, 3). This means that the microcavity is in the vacuum state and the QD is in the lower dressed state |−〉, corresponding to ρ(∞) = |0〉 〈0| ⊗ ρb(∞). In fact, this condition γ = 0 can be realized by technically engineering the microcavity into the PBG material. As mention before, by properly engineering the system topology, the QD coupled to the microcavity can experiences a discontinuity of the LDOS. It strongly influences the dynamics of QD and results in new emission effects, i.e. |D(ωλ)|2 = 0 for ωλ < ωb and |D(ωλ)|2 = 1 for ωλ > ωb. In this way, the spontaneous emissions occurring at frequencies ω and ωL are selectively eliminated, i.e., both the spontaneous emission rates γ and γ0 are zero. Only the spontaneous emission occurring at the frequency ω+ = ωL + 2Ω̄ is allowed to happen with the spontaneous emission rate γ+. In our scheme, a nonlinear coherent state of the MR mode is generated based on the presence of a dark state. Additionally, if the PBG material were to be moved such that none of the spontaneous transitions are eliminated, the steady state of the system can not evolve into such a dark state, then the scheme for generation of the target state may be the other project which is out of our interest.

When taking the couplings of the microcavity with the QD and the weak laser field into account, i.e., V1 ≠ 0, V′2 ≠ 0, we devote to finding the solution which simultaneously satisfy the equations −i[V1 + V′2, ρ] = 0 and ℒρ = 0. Because the solution of the latter one is that in which the QD-cavity subsystem is in the state |0〉 as discussed above. Therefore the steady state of the whole system can be described by |Ψ〉 = |0, φb〉, (ρ = |0, φb〉 〈0, φb|), where |φb〉 is the motional state of the mechanical mode. When the total system is in such a state, the QD-microcavity system stops to fluoresce and the motional state |φb〉 can be obtained. Therefore, it can be said that the system is in a motional dark state [21], which has various types. This target state can be obtained from equation

(V1+V2)|0,φb=0,
which means the phonon state |φb〉 satisfying
[f1(bb)b+χf22(bb)]|φb=0.
We call the eigenstates of Eq. (33) with the boson annihilation operator b and nonlinear functions of number operator = bb as NCSs [21]. At this point, we can see that the state |φb〉 is a NCS of the MR which is obtained by using the dark state properties of the system.

In order to explore the properties of |φb〉, we expand the state |φb〉 as

|φb=n=0cn|n,
in which cn is the probability amplitude of the Fock state |n〉. In the Fock representation, we use the definitions of fi(bb), (i = 1, 2) in Eq. (15) and Eq. (17) to expand Eq. (33). After some straightforward calculations, we find that the coefficients of the Fock states satisfy a recursion relation as
n+1χcn[Ln(0)(η2)]2cn+1Ln(1)(η2)=0,
where Ln(m)(η2) with the definition Ln(m)(x)=k=0(1)k(m+n)!(nk)!(k+m)!xkk! are associated Laguerre polynomials. In the numerical calculation of this paper, we choose the Lamb-Dicke parameter η0 such that Ln(0)(η2)=0 for n = N. For a given positive integer N, the Hilbert space can be confined within N phonon number states if one sets the transition coefficients to the N + 1 phonon number states cn+1 zero. This recursion relation ensures that all coefficients cn vanish for n > N. Thus, there will be none excitations from N phonon states to N + 1 phonon states.

So far, by designing a suitable coherent Hamiltonian and the interaction between the phonon bath and the coupled QD-cavity system, the mechanical mode can be prepared to the NCS |φb〉 which is a non-classical motional dark state. It should be highlighted here that the prepared state |φb〉 of the mechanical mode is completely decoupled from the QD and the cavity. This can be understood in the polaron frame described by Eq. (7), because the whole system can be described by |Ψ〉 = |0, φb〉, where the state |φb〉 of mechanical mode is decoupled from the QD and the cavity. However, when going back from this polaron frame, we obtain

|Ψ˜=eS|0,φb=|0,φb=|Ψ,
where the operator S is defined below Eq. (7), i.e. S = ηaa(bb). |0〉 = |−, 0〉 is the dark state of the reduced QD-cavity subsystem, in which the cavity is in the vacuum state and the QD is in the lower dressed state |−〉. As a result, |Ψ〉 = |0, φb〉 keeps the same form under the inverse polaron transformation. Therefore, the prepared NCS of the MR is completely decoupled from the QD and the cavity.

Recently, when a two-level system and a mechanical mode are both strongly coupled to a cavity, the preparation of the mechanical modes in a non-Gaussian dark state has also been studied [70]. Employing additional auxiliary cavity mode to construct two Liouvillians, an approximated entangled state are prepared in the weak coupling limit, i.e., ζ = (gq/gm)2 ≤ 1, where gq and gm are coupling strengthes of the cavity with a two-level system and a mechanical mode respectively. When ignoring the spontaneous emission of the two-level system and measuring the two-level system, the mechanical oscillator can be prepared in a superposition of Fock states with a fidelity of F ≈ 0.83. However, it should be emphasized that the literature [70] is distinct from this paper. In this paper, when properly tuning the frequency of the weak driving field, two-photon blockade phenomenon occurs. The QD-cavity subsystem can evolve into a dark state due to the damping of the microcavity and the elimination of the decay rate of the QD at selected frequencies in the photonic band-gap (PBG) material. By exploiting the interaction between the phonon bath and the coupled QD-cavity system, a suitable coherent Hamiltonian is designed so that the mechanical mode can naturally be prepared into a NCS. The prepared NCS of the MR in this paper is a definite pure state. Besides, the prepared NCS is completely decoupled from the QD and the cavity. The intrinsic non-linearity and non-classical behavior of the state will be analysed via phonon number statistics, second-order correlation function, and Wigner function.

3.1. Phonon number statistics

The phonon statistics of a MR are the “finger-print” of its quantum state, from which a number of useful measures of non-classicality may be inferred. The phonon number distribution (PND) of the dark states is complicated by analytical ways with the correlated polynomials. In this paper, we obtain the results by numerically solving the equation involving Laguerre polynomial. For simplicity, we factitiously choose the parameter η = g2/ωm = 0.305463 to satisfy the equation

Ln(0)(η2)=0.
In this case, we have numerically checked that for N > 15, the transition coefficients to the N + 1 phonon number states vanish.

We should note that η = g2/ωm = 0.305463 means the system approaches the single-photon strong coupling regime. If the single-photon optomechanical coupling strength g0 is an appreciable fraction of the mechanical resonance frequency ωm and the combined nonlinearity parameter satisfies the relation g02/κωm1, the system is in the single-photon ultra-strong coupling regime characterized by cavity photon blockade [71, 72]. This challenging idea has inspired emerging attentions from experimental field. The nonlinearity of the phonon mode of the MR originating from the strong coupling between the microcavity and the MR has been discussed based on capacitive [73] or magnetic [74] coupling in the microwave-regime optomechanical devices. By enhancing the effect of a single-photon in the optical cavity, the strong-coupling regime is reached with the auxiliary of the squeezing of a cavity mode [4]. A single-photon ultra-strong coupling regime is proposed to be reached in an optomechanical system [75]. For a system with a quantum dot strongly coupled to a photonic-crystal nanocavity, J. Vučković et. al have studied non-classical higher-order photon correlations [76]. The strong coupling of a single two-level solid-state system with a photon is realized with a single quantum dot in a semiconductor microcavity [77]. Here, we supply some main parameters to implement our scheme with ωc/2π = 195THz and ωm/2π = 3.68GHz, κ/2π = 150MHz, g1 = 5κ, which are similar to that in the literature [10]. The condition η = g2/ωm = 0.3 indicates that a value of g2/2π = 1.1GHz is needed to be achieved. At present, the optomechanical coupling strength g2 has realized g2/2π = 11.5MHz, and is predicted to approach g2/2π = 26MHz [65]. It is still inadequate for realizing our proposal. However, very recently, it has been proposed that applying a large-amplitude, strongly detuned mechanical parametric drive to amplify mechanical zero-point fluctuations, the radiation pressure interaction is enhanced and thus the overall coupling strength is dramatically increased, allowing us to completely enter the quantum-enabled, strong-coupling regime [78]. The strong radiation pressure forces from squeezed light on a mechanical oscillator is observed [79]. These inspiring and promising progresses enable us to believe that our scheme can be achieved in the near future.

In Eq. (37), the fact cn+1 = 0 for n > N = 15 indicates that the MR is trapped into the Fock superposition state via the cavity damping and selectively elimination the spontaneous emission γ in the PBG material. The specific expression of cn is

cn=χn![j=0n1Lj1(0)(η2)]2j=0n1Lj1(1)(η2)c0,
which can be controlled by adjusting χ. The PND is defined as
P(n)=Tr(ρ|nn|),
where |n〉 is the eigenstate of the phonon number operator bb. The probability of finding n phonons in the field is given by P(n) = |cn|2. Note that for the NCS, the probability of the phonon number depends on the nonlinearity functions and varies with parameter αp/g1. In figure 4, we show the PND of the MR as a function of n for the different values of αp/g1. From this figure it is seen that the probability is nearly close to zero for n ≥ 7. As apparently shown in the figure 4, the peak of Pn depends on the parameter αp/g1, and which position moves forward to the larger number distribution with increasing of the parameter αp/g1, meanwhile value of the peak is reduced. With the increasing of the parameter αp/g1, the mean phonon number in the NCS may increase, which means that multi-phonon processes are enhanced.

 

Fig. 4 The PND Pn of the MR state in the steady-state regime versus n for different values of χ: the blue corresponds to αp/g1 = 0.11; the red corresponds to αp/g1 = 0.15. The other parameters are as follows: N = 15, Δa = 0, η = 0.305463.

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3.2. Second-order correlation function(SOCF)

In order to extract more information about the NCS of the MR, we calculate the corresponding second-order correlation function(SOCF) which is correlated with the coherence and statistical properties a field. The normalized SOCF of the phonon mode of the MR in the steady state can be defined as [80]

g(2)(0)=b2b2sbbs.
where the superscript s describes the steady state of the field and b(b) is the creation (annihilation) operator of the MR. This function can give the influences of detecting one phonon of the source on the result of detecting another one. The SOCF g(2)(0) can be applied to distinguish whether the statistical properties of the field is super-Poissonian (g(2)(0) > 1), Poissonian (g(2)(0) = 1), or sub-Poissonian (g(2)(0) < 1). The sub-Poissonian statistics is usually correlated to non-classical state [81, 82].

In order to get the insight into the non-classical properties of the phonon, we numerically calculate the SOCF in the Fock state representation with the following expression

g(2)(0)=n=0n(n1)|cn|2(n=0n|cn|2)2.
As a result, we demonstrate the SOCF of the phonon as the function of the parameter αp/g1 in figure 5. It is clearly shown from figure 5 that the value of SOCF g(2)(0) decreases with the increasing of αp/g1 and approaches the vicinity of g(2)(0) < 1 when the parameter αp/g1 approximately takes the value 0.1225. The fact g(2)(0) < 1 indicates the sub-Poissonian statistics and the non-classicality of the generated phonon state. It also implies that the NCS of the MR |φb〉 exhibits the sub-Poissonian behaviour and non-classical effect.

 

Fig. 5 The second-order correlation function Pn of the MR state in the steady-state regime versus αp/g1. The other parameters are the same as those in figure 4.

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3.3. Wigner function

The Wigner function (WF) [83] was first introduced by Wigner in 1932 to calculate quantum correction to a classic distribution function of a quantum-mechanical system. It is a quasi-probability distribution which fully describes the states of a quantum system in phase space, (either the position-momentum space for an harmonic oscillator or, equivalently, the space spanned by two orthogonal quadratures of the electromagnetic fields for a single-mode state of light), in the same fashion as a probability distribution (non-negative by definition) characterizes a classical system. It has now become a useful tool to study the non-classical properties of quantum states, because the partial negativity of the WF [83,84] is indeed a good indication of the highly non-classical character of the state. In order to further unveil the non-classicality of the NCS of phonon, in this subsection, we supply the expression of WF for the phonon in Fock state representation and numerically illustrate the corresponding property.

For a single-mode system, the WF of the density operator ρ in Fock state representation |n〉 is expressed as

W(β)=2πn=0(1)nn|D(β)ρD(β)|n,
in which ρ is the reduced density operator of the MR and D(β) = exp(βbβ*b) is the displacement operator. The WF of the NCS |φb〉 is depicted in figure 6 in phase space, where and Ŷ correspond to the position and the momentum respectively. The distribution is suppressed along quadrature component and presents crescent shape. According to Hudson’s theorem [85], the presence of negative regions in the Wigner distribution is an authentic sign of quantum character. Obviously, there is a negative region in figure 6, which reveals the strong non-classicality and genuine quantum behavior.

 

Fig. 6 The WF of the phonon in the steady-state regime with αp/g1 = 0.15. The other parameters are the same as those in Figure 4.

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As mentioned at the beginning of this subsection, the partial negativity of the WF [83, 84] is a good indication of the highly non-classical character of a quantum state. In order to quantitatively explore the non-classicality of the state, we study the absolute value PW of the total negative probability of the WF, which can be defined by [84, 86]

PW=|ϖW(x,y)dxdy|,
in which ϖ is the negative region of WF distribution. Since the negative region W(x, y) of WF has been already obtained. Then, the value of PW can be numerically calculated according to the above definition. One should note in Eq. (43) that only the negative value of W(x, y) is integrated. During the calculation, the value of PW is set to be zero when W(x, y) > 0.

We illustrate PW of the total negative probability of the WF as a function of the parameter αp/g1 in figure 7. It can be seen from figure 7 that with the increasing of αp/g1, PW takes a definite value when the parameter αp/g1 approaches the region αp/g1 > 0.09. The peak of PW occurs at αp/g1 ≈ 0.145. Whereas, as found from figure 5 that g(2)(0) takes a smaller value than 1 in the region αp/g1 ≥ 0.1225. In this region 0.09 < αp/g1 < 0.1225, g(2)(0) takes a larger value than 1, being unable to tell whether or not the non-classical effect appears. But in the same region PW takes a definite value, which is the indication of non-classicality. The region for g(2)(0) < 1 is narrower than that for PW > 0. In other words, the WF of the phonon state has a negative region and the state possesses the non-classicality, but it does not show the sub-poissonian phonon statistics [87]. Therefore, we emphasize here that for the prepared NCS of the phonon mode in this paper, the absolute value PW may be a better choice for quantifying the non-classicality than the SOCF.

 

Fig. 7 The absolute value PW of the total negative probability of the Wigner function as a function of the parameter αp/g1. The other parameters are the same as those in figure 4.

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Up to now, we have prepared the NCS of the phonon mode of MR by using the dark state of the system. Both the facts that g(2)(0) < 1 and the WF approaches a negative region are convincing indications of the sub-Poissonian behaviour and non-classicality of the NCS. However, we should point out here that in all the calculations, the damping of the MR is neglected. Because for a MR with high quality factor, the mechanical damping γm could be much smaller than the cavity dissipation rate κ. The mechanical damping can be neglected during the time needed to achieve the steady states. As mentioned before, in our scheme, the four states concerned in Eq. (19) are established in the subspace of the reduced QD-cavity subsystem, which involves the strong driving field ωL. The resonant frequency of the cavity ωc is at least 103 times the frequency of ωm [8–10, 50, 62–66]. In this situation, the energy interval between the subspaces of the reduced QD-cavity system is much larger than the energy interval of the MR Fock space. Therefore, the four states of the QD-cavity subsystem and 15 Fock states of the MR are valid to describe the considered system. Here, we should mention that the increasing of ωm corresponds to a larger quality factor, which is estimated to 1.1 × 106 when ωm = 2π × 5.3GHz at cryogenic temperature [50]. A larger ωm than the decay of the cavity κ means the appearance of the sideband excitations. Then, the sideband cooling method can be employed. For a given cavity matched with the QD, the increasing of ωm may dissatisfy the condition g02/κωm1 thus lead to difficult achievement of strong coupling regime, where g0 is the optomechanical coupling strength. In this case, as discussed before, it has been proposed recently that applying a large-amplitude, strongly detuned mechanical parametric drive to amplify mechanical zero-point fluctuations, the radiation pressure interaction is enhanced and thus the overall coupling strength is dramatically increased, allowing one to completely enter the quantum-enabled, strong-coupling regime [78]. The strong radiation pressure forces from squeezed light on a mechanical oscillator is observed [79]. These inspiring and promising progresses enable us to believe that our scheme can be achieved in the near future.

As is known that non-Gaussian states might be more appropriate than Gaussian state in the sense of the robustness in the applications of teleportation and the cloning [88–90] of quantum states. The nonlinear coherent state generated in the solid-state optomechanical system may benefit for the optomechanical quantum information processing with photons and phonons [36].

4. Conclusion

In this paper, we have proposed a scheme to prepare a NCS for the MR by using the dark state of the optomechanical system, in which a two-level QD located inside a microscopic cavity is engineered in a PBG material. The effective master equation of the optomechanical system is obtained in the polaron frame. Two-photon transition can be realized in the dressed states of the QD-microcavity subsystem via adjusting the discrepancy of the frequencies between the micro-cavity and the weak laser field. By exploiting the interaction between the phonon bath and the coupled QD-cavity system, a suitable coherent Hamiltonian is designed so that the mechanical mode can naturally be prepared into a NCS, which is a definite pure state and is completely decoupled from the QD and the cavity. The NCS of the phonon exhibits strong non-classicality since its SOCF shows the sub-Poissonian distribution and its Wigner function approaches the negative region. Our proposal provides a way to prepare a NCS with nonclassicality in an optomechanical crystal system. We hope this could stimulate discussions about the applications of non-Gaussian non-classical states in the quantum information processing [36] and quantum teleportation [91] in the optomechanical systems. The NCS could contribute to the application in optimizing quantum information in the near future, which is important in binary (or multibinary) communication [27].

Acknowledgments

This work was supported by the National Natural Science Foundation of China (NSFC) (Grants No. 61275123, No.11474119, No.11404103 and No. 11304024) and the National Basic Research Program of China (Grant No. 2012CB921602).

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H. Fakhri and A. Hashemi, “Nonclassical properties of the q-coherent and q-cat states of the Biedenharn-Macfarlane q oscillator with q>1,” Phys. Rev. A 93, 013802 (2016).
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R. Riedinger, S. Hong, R. A. Norte, J. A. Slater, J. Shang, A. G. Krause, V. Anant, M. Aspelmeyer, and S. Gröblacher, “Non-classical correlations between single photons and phonons from a mechanical oscillator,” Nature 530, 313–316 (2016).
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P. D. Nation, J. Suh, and M. P. Blencowe, “Ultrastrong optomechanics incorporating the dynamical Casimir effect,” Phys. Rev. A 93, 022510 (2016).
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M. A. Lemonde, N. Didier, and A. A. Clerk, “Enhanced nonlinear interactions in quantum optomechanics via mechanical amplification,” Nat. Commun. 7, 11338 (2016).
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2015 (9)

B. M. Ann, Y. Song, J. Kim, D. Yang, and K. An, “Correction for the detector-dead-time effect on the second-order correlation of stationary sub-Poissonian light in a two-detector configuration,” Phys. Rev. A 92, 023830 (2015).
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Y. H. Zhou, H. Z. Shen, and X. X. Yi, “Unconventional photon blockade with second-order nonlinearity,” Phys. Rev. A 92, 023838 (2015).
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J. M. Pirkkalainen, S. U. Cho, F. Massel, J. Tuorila, T. T. Heikkilä, P. J. Hakonen, and M. A. Sillanpää, “Cavity optomechanics mediated by a quantum two-level system,” Nat. Commun. 6, 6981 (2015).
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G. Via, G. Kirchmair, and O. Romero-Isart, “Strong single-photon coupling in superconducting quantum magnetomechanics,” Phys. Rev. Lett. 114, 143602 (2015).
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R. Leijssen and E. Verhagen, “Strong optomechanical interactions in a sliced photonic crystal nanobeam,” Sci. Rep. 5, 15974 (2015).
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F. Lecocq, J. D. Teufel, J. Aumentado, and R. W. Simmonds, “Resolving the vacuum fluctuations of an optomechanical system using an artificial atom,” Nature Phys. 11, 635–639 (2015).
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W. W. Deng, G. X. Li, and H. Qin, “Enhancement of the two-photon blockade in a strong-coupling qubit-cavity system,” Phys. Rev. A 91, 043831 (2015).
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S. Dey, “Q-deformed noncommutative cat states and their nonclassical properties,” Phys. Rev. D 91, 044024 (2015).
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X. Y. Lü, Y. Wu, J. R. Johansson, H. Jing, J. Zhang, and F. Nori, “Squeezed optomechanics with phase-matched amplification and dissipation,” Phys. Rev. Lett. 114, 093602 (2015).
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2014 (6)

W. J. Gu, G. X. Li, S. P. Wu, and Y. P. Yang, “Generation of non-classical states of mirror motion in the single-photon strong-coupling regime,” Opt. Express 22, 18254–18267 (2014).
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S. Carlig and M. A. Macovei, “Long-time correlated quantum dynamics of phonon cooling,” Phys. Rev. A 90, 013817 (2014).
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A. C. Pflanzer, O. Romero-Isart, and J. I. Cirac, “Optomechanics assisted by a qubit: from dissipative state preparation to many-partite systems,” Phys. Rev. A 88, 033804 (2014).
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A. Rundquist, M. Bajcsy, A. Majumdar, T. Sarmiento, K. Fischer, K. G. Lagoudakis, S. Buckley, A. Y. Piggott, and J. Vučković, “Nonclassical higher-order photon correlations with a quantum dot strongly coupled to a photonic-crystal nanocavity,” Phys. Rev. A 90, 023846 (2014).
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M. Aspelmeyer, T. J. Kippenberg, and F. Marquardt, “Cavity optomechanics,” Rev. Mod. Phys. 86, 1391–1452 (2014).
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J. Restrepo, C. Ciuti, and I. Favero, “Single-polariton optomechanics,” Phys. Rev. Lett. 112, 013601 (2014).
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2013 (8)

X. W. Xu, H. Wang, J. Zhang, and Y. X. Liu, “Engineering of nonclassical motional states in optomechanical systems,” Phys. Rev. A 88, 063819 (2013).
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A. Miranowicz, M. Paprzycka, Y. X. Liu, J. Bajer, and F. Nori, “Two-photon and three-photon blockades in driven nonlinear systems,” Phys. Rev. A 87, 023809 (2013).
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X. W. Xu, Y. J. Zhao, and Y. X. Liu, “Entangled-state engineering of vibrational modes in a multimembrane optomechanical system,” Phys. Rev. A 88, 022325 (2013).
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A. Kronwald, F. Marquardt, and A. A. Clerk, “Arbitrarily large steady-state bosonic squeezing via dissipation,” Phys. Rev. A 88, 063833 (2013).
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T. Ramos, V. Sudhir, K. Stannigel, P. Zoller, and T. J. Kippenberg, “Nonlinear quantum optomechanics via individual intrinsic two-level defects,” Phys. Rev. Lett. 110, 193602 (2013).
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W. J. Gu, G. X. Li, and Y. P. Yang, “Generation of squeezed states in a movable mirror via dissipative optomechanical coupling,” Phys. Rev. A 88, 013835 (2013).
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Y. C. Liu, Y. F. Xiao, X. S. Luan, and C. W. Wong, “Dynamic dissipative cooling of a mechanical resonator in strong coupling optomechanics,” Phys. Rev. Lett. 110, 153606 (2013).
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W. J. Gu and G. X. Li, “Quantum interference effects on ground-state optomechanical cooling,” Phys. Rev. A 87, 025804 (2013).
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2012 (7)

E. Verhagen, S. Deléglise, S. Weis, A. Schliesser, and T. J. Kippenberg, “Quantum-coherent coupling of a mechanical oscillator to an optical cavity mode,” Nature 482, 63–67 (2012).
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D. W. C. Brooks, T. Botter, S. Schreppler, T. P. Purdy, N. Brahms, and D. M. Stamper-Kurn, “Non-classical light generated by quantum-noise-driven cavity optomechanics,” Nature 488, 476–480 (2012).
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A. H. Safavi-Naeini, J. Chan, J. T. Hill, T. P. M. Alegre, A. Krause, and O. Painter, “Observation of quantum motion of a nanomechanical resonator,” Phys. Rev. Lett. 108, 033602 (2012).
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S. Ya. Kilin and A. B. Mikhalychev, “Single-atom laser generates nonlinear coherent states,” Phys. Rev. A 85, 063817 (2012).
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S. Dey and A. Fring, “Squeezed coherent states for noncommutative spaces with minimal length uncertainty relations,” Phys. Rev. D 86, 064038 (2012).
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K. Stannigel, P. Komar, S. J. M. Habraken, S. D. Bennett, M. D. Lukin, P. Zoller, and P. Rabl, “Optomechanical quantum information processing with photons and phonons,” Phys. Rev. Lett. 109, 013603 (2012).
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R. Jacob, S. Winnerl, M. Fehrenbacher, J. Bhattacharyya, H. Schneider, M. T. Wenzel, H. G. von Ribbeck, L. M. Eng, P. Atkinson, O. G. Schmidt, and M. Helm, “Intersublevel spectroscopy on single InAs-quantum dots by terahertz near-field microscopy,” Nano Lett. 12, 4336–4340 (2012).
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2011 (9)

C. Roy and S. Hughes, “Phonon-dressed Mollow triplet in the regime of cavity quantum electrodynamics: excitation-induced dephasing and nonperturbative cavity feeding effects,” Phys. Rev. Lett. 106, 247403 (2011).
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C. Lang, D. Bozyigit, C. Eichler, L. Steffen, J. M. Fink, A. A. Abdumalikov, M. Baur, S. Filipp, M. P. da Silva, A. Blais, and A. Wallraff, “Observation of resonant photon blockade at microwave frequencies using correlation function measurements,” Phys. Rev. Lett. 106, 243601 (2011).
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A. J. Hoffman, S. J. Srinivasan, S. Schmidt, L. Spietz, J. Aumentado, H. E. Türeci, and A. A. Houck, “Dispersive photon blockade in a superconducting circuit,” Phys. Rev. Lett. 107, 053602 (2011).
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J. Chan, T. P. M. Alegre, A. H. Safavi-Naeini, J. T. Hill, A. Krause, S. Gröblacher, M. Aspelmeyer, and O. Painter, “Laser cooling of a nanomechanical oscillator into its quantum ground state,” Nature 478, 89–92 (2011).
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J. Q. Liao and C. K. Law, “Parametric generation of quadrature squeezing of mirrors in cavity optomechanics,” Phys. Rev. A 83, 033820 (2011).
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A. H. Safavi-Naeini, T. P. M. Alegre, J. Chan, M. Eichenfield, M. Winger, Q. Lin, J. T. Hill, D. E. Chang, and O. Painter, “Electromagnetically induced transparency and slow light with optomechanics,” Nature 472, 69–73 (2011).
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J. D. Teufel, T. Donner, D. Li, J. W. Harlow, M. S. Allman, K. Cicak, A. J. Sirois, J. D. Whittaker, K. W. Lehnert, and R. W. Simmonds, “Sideband cooling of micromechanical motion to the quantum ground state,” Nature 475, 359–363 (2011).
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2010 (5)

A. D. O’Connell, M. Hofheinz, M. Ansmann, R. C. Bialczak, M. Lenander, E. Lucero, M. Neeley, D. Sank, H. Wang, M. Weides, J. Wenner, J. M. Martinis, and A. N. Cleland, “Quantum ground state and single-phonon control of a mechanical resonator,” Nature 464, 697–703 (2010).
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M. G. Genoni and M. G. A. Paris, “Quantifying non-Gaussianity for quantum information,” Phys. Rev. A,  82, 052341 (2010).
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D. Englund, A. Majumdar, A. Faraon, M. Toishi, N. Stoltz, P. Petroff, and J. Vuckovic, “Resonant excitation of a quantum dot strongly coupled to a photonic crystal nanocavity,” Phys. Rev. Lett. 104, 073904 (2010).
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J. Niset, J. Fiurasek, and N. J. Cerf, “No-go theorem for Gaussian quantum error correction,” Phys. Rev. Lett. 102, 120501 (2009).
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X. Ma and S. John, “Switching dynamics and ultrafast inversion control of quantum dots for on-chip optical information processing,” Phys. Rev. A 80, 063810 (2009).
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A. P. Lund, T. C. Ralph, and H. L. Haselgrove, “Fault-tolerant linear optical quantum computing with small-amplitude coherent states,” Phys. Rev. Lett. 100, 030503 (2008).
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M. B. Harouni, R. Roknizadeh, and M. H. Naderi, “Nonlinear coherent state of an exciton in a wide quantum dot,” J. Phys. B: At. Mol. Opt. Phys. 41, 225501 (2008).
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M. Li, H. X. Tang, and M. L. Roukes, “Ultra-sensitive NEMS-based cantilevers for sensing, scanned probe and very high-frequency applications,” Nat. Nanotecnol. 2, 114–120 (2007).
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F. DellAnno, S. De Siena, L. Albano, and F. Illuminati, “Continuous-variable quantum teleportation with non-Gaussian resources,” Phys. Rev. A 76, 022301 (2007).
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S. Olivares and M. G. A. Paris, “De-Gaussification by inconclusive photon subtraction,” Laser Phys. 16, 1533–1550 (2006).
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N. J. Cerf, O. Kruger, P. Navez, R. F. Werner, and M. M. Wolf, “Non-Gaussian cloning of quantum coherent states is optimal,” Phys. Rev. Lett. 95, 070501 (2005).
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Figures (7)

Fig. 1
Fig. 1 Schematic of the considered system. A photonic crystal nanobeam is clamped at both ends. The optomechanical microcavity with the frequency ωc can be coupled with the MR of frequency ωm via a combination of radiation pressure and photostriction [49]. The MR can travel fast and alternating expansions and contractions along the x axis, which we refer to as a “breathing” mode [50]. A QD (the small red dot in Fig) driven by a strong laser is within the photonic crystal microcavity.
Fig. 2
Fig. 2 Schematic representation of the LDOS and the relative position of the relevant frequencies considered in our paper. ωb, ωc andωL are the photonic density-of-states bandedge frequency, the cavity frequency and the coherent driving frequency respectively. The vertical arrows represent the emission lines of the Mollow spectrum of the QD resonance fluorescence which central band occurs at the frequency ωL, and the side bands components at ωL ± 2Ω̄.
Fig. 3
Fig. 3 Schematic of the four-level model. The two-photon transition from the ground state |0〉 to the higher excited state |3〉 is resonant.
Fig. 4
Fig. 4 The PND Pn of the MR state in the steady-state regime versus n for different values of χ: the blue corresponds to αp/g1 = 0.11; the red corresponds to αp/g1 = 0.15. The other parameters are as follows: N = 15, Δ a = 0, η = 0.305463.
Fig. 5
Fig. 5 The second-order correlation function Pn of the MR state in the steady-state regime versus αp/g1. The other parameters are the same as those in figure 4.
Fig. 6
Fig. 6 The WF of the phonon in the steady-state regime with αp/g1 = 0.15. The other parameters are the same as those in Figure 4.
Fig. 7
Fig. 7 The absolute value PW of the total negative probability of the Wigner function as a function of the parameter αp/g1. The other parameters are the same as those in figure 4.

Equations (43)

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H = H 0 + H I ,
H 0 = Δ a 2 σ z + Ω ( σ + + σ ) + Δ c a a + ω m b b + λ Δ λ a λ a λ ,
H I = [ α p a e i Δ p t + g 1 a σ + λ g λ ( ω λ ) a λ σ + h . c . ] + g 2 a a ( b + b ) ,
| + = s | g + c | e , | = c | g s | e ,
H 0 = Ω ¯ R 3 + Δ c a a + ω m b b + λ Δ λ a λ a λ .
H I = { [ g 1 a + λ g λ ( ω λ ) a λ ] ( s c R 3 + c 2 R + s 2 R + ) + h . c . } + α p ( a e i Δ p t + a e i Δ p t ) + g 2 a a ( b + b ) ,
H = e S ( H 0 + H I ) e S ,
H = H 0 + H I 1 + H I 2 + H I 3
H 0 = Ω ¯ R 3 + Δ c a a + ω m b b + λ Δ λ a λ a λ ,
H I 1 = g 1 a e η ( b b ) ( s c R 3 + c 2 R + s 2 R + ) ,
H I 2 = α p a e η ( b b ) e i Δ p t + h . c . ,
H I 3 = λ g λ ( ω λ ) a λ ( s c R 3 + c 2 R + s 2 R + ) + h . c .
e η ( b b ) = e η 2 2 m , n = 0 η m ( η ) n m ! n ! b m b n .
H ˜ I 1 ( t ) = g 1 [ c 2 a R + η s 2 a f 1 ( b b ) b R + e i δ m t + h . c . ] ,
f 1 ( b b ) = n = 0 ( 1 ) n + 1 η 2 n n ! ( n + 1 ) ! b n b n .
H ˜ I 2 ( t ) = α p [ a f 2 ( b b ) e i δ cp t + h . c . ] ,
f 2 ( b b ) = n = 0 ( 1 ) n η 2 n n ! n ! b n b n .
H ˜ I 3 = λ g λ ( ω λ ) a λ [ s c R 3 e i Δ λ t + c 2 R + e i ( Δ λ 2 Ω ¯ t ) s 2 R + e i ( Δ λ + 2 Ω ¯ ) t ] + h . c .
| 0 = | , 0 , | 1 = | , 1 | + , 0 2 , | 2 = | , 1 + | + , 0 2 , | 3 = | , 2 | + , 1 2 .
H J = i = 0 3 E i | i i | .
a = 1 2 ( | 1 + | 2 ) 0 | + | 3 ( 2 + 1 2 1 | + 2 1 2 2 | ) , R + = 1 2 | 0 ( 1 | 2 | ) 1 2 ( | 1 + | 2 ) 3 | .
V I = V 1 + V 2 ,
V 1 = Ω e [ | 3 0 | f 1 ( b b ) b + h . c . ] ,
V 2 = α p [ 2 2 | 1 0 | e i Δ 1 t + 2 2 | 2 0 | e i Δ 2 t + 2 + 1 2 | 3 1 | e i Δ 3 t + 2 1 2 | 3 2 | e i Δ 4 t ] f 2 ( b b ) + h . c .
V 2 = i V 2 ( t ) V 2 ( t ) d t .
V 2 = Ω e [ χ f 2 2 ( b b ) | 3 0 | + h . c . ]
ρ t = 0 t d t Tr R { [ H ˜ int ( t ) , [ H ˜ int ( t ) , ρ T ( t ) ] ] } ,
λ ρ = γ 0 𝒟 [ R 3 ] + γ 𝒟 [ R + ] + γ + 𝒟 [ R + ] ,
c ρ = κ 𝒟 [ a e η ( b b ) ] .
d d t ρ = i [ V 1 + V 2 , ρ ] + λ ρ + c ρ ,
λ ρ = j = 1 2 { γ 2 𝒟 [ | j 0 | ] + γ + 2 𝒟 [ | 0 j | ] + γ 4 𝒟 [ | 3 j | ] + γ + 4 𝒟 [ | j 3 | ] + γ 0 { 𝒟 [ | 0 0 | ] + 𝒟 [ | 1 2 | ] + 𝒟 [ | 2 1 | ] } , c ρ = κ e η 2 4 { j = 1 2 m , n , k = 0 ϒ b m b n | 0 j | ρ | j 0 | b k b m + k n + j = 1 2 m , n , k = 0 B j ϒ b m b n | j 3 | ρ | 3 j | b k b m + k n } κ e η 2 4 j = 1 2 ( 1 + B j ) ( | j j | ρ + ρ | j j | ) ,
( V 1 + V 2 ) | 0 , φ b = 0 ,
[ f 1 ( b b ) b + χ f 2 2 ( b b ) ] | φ b = 0 .
| φ b = n = 0 c n | n ,
n + 1 χ c n [ L n ( 0 ) ( η 2 ) ] 2 c n + 1 L n ( 1 ) ( η 2 ) = 0 ,
| Ψ ˜ = e S | 0 , φ b = | 0 , φ b = | Ψ ,
L n ( 0 ) ( η 2 ) = 0 .
c n = χ n ! [ j = 0 n 1 L j 1 ( 0 ) ( η 2 ) ] 2 j = 0 n 1 L j 1 ( 1 ) ( η 2 ) c 0 ,
P ( n ) = Tr ( ρ | n n | ) ,
g ( 2 ) ( 0 ) = b 2 b 2 s b b s .
g ( 2 ) ( 0 ) = n = 0 n ( n 1 ) | c n | 2 ( n = 0 n | c n | 2 ) 2 .
W ( β ) = 2 π n = 0 ( 1 ) n n | D ( β ) ρ D ( β ) | n ,
P W = | ϖ W ( x , y ) d x d y | ,

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