Nonlinear coupling between a nitrogen-vacancy-center ensemble and a superconducting qubit

Qiong Chen; Jun Wen; W. L. Yang; M. Feng; Jiangfeng Du

doi:10.1364/OE.23.001615

1. Introduction

In quantum optics, quantum nonlinear effects are indispensable for investigating and controlling the dynamics of harmonic oscillators [1]. Especially, the giant nonlinearities would open up many potential applications, such as generating nonclassical field states [2], squeezing and parametric amplification, [1, 3–5], which have been regarded as important resources for universally processing quantum information with continuous variables [6]. Therefore, it is highly desirable to develop new quantum hybrid systems to realize strong nonlinear couplings [5–11], especially in the solid-state systems [5–15].

Recently, the strong coupling regime has been experimentally implemented between super-conducting qubit (SQ) and transmission line resonator (TLR) system [16–18], and many important quantum information processes (QIPs) [19], including the coupling of Josephson qubits by the TLR as a quantum data bus [20] and the preparation of the TLR Fock states [18], have been demonstrated. Based on these experimental advances, there are various of interests in investigating nonlinear interactions in circuit QED systems [21–25]. For example, a phase-preserving amplifier based on the nonlinear interaction was proposed and realized to improve its performance in the regime near the quantum limit [23]. By using a single artificial multilevel Cooper pair box molecule, one may achieve giant self-Kerr effect at microwave frequencies [12]. Additionally, there are promising efforts to realize squeezed states of microwaves by using three-wave mixing coupling in circuit QED [4, 5].

The nitrogen-vacancy defect center (NV) in diamond is of particular interest because it offers unprecedented tunability and scalability, which are easily integrated with other quantum systems [26, 27]. Moreover, it is optically controllable and has extremely long nuclear and electronic decoherence time [28]. Due to the enhanced magnetic-dipole coupling through the collective excitation of the spins in NV ensemble (NVE) involving large numbers of NV centers with N = 10¹¹ ∼ 10¹³ [31], much effort has been devoted recently to combine the NVE with circuit quantum electrodynamics (QED) both experimentally and theoretically [31–38]. There are many significant progresses in simulating quantum optics phenomena and implementing quantum information processing, such as quantum memories and continuous variable entanglement [34–36, 38]. However, no nonlinear coupling in circuit QED has been considered yet in the combined NVE-circuit QED hybrid systems.

Here we focus on the weak quadratic coupling between the SQ and TLR, and realize a new kind of nonlinear interaction between spin ensemble and SQ in the hybrid system consisting of a NVE, a SQ and a TLR. By virtual excitation of microwave photons, which acts as a quantum bus, an effective nonlinear coupling between the NVE and the SQ is generated in the dispersive limit. More intriguingly, we consider the total spin number of NV centers as N = 10¹¹ ∼ 10¹³ [31, 32], where the nonlinear NVE-SQ coupling actually benefits from the collective enhancement. Physically, as long as the ensemble has only few excited spins, the excitation behaves like a harmonic oscillator degree of freedom in the low-excitation regime. Therefore, a pair of bosons in the NVE is excited when the SQ decays from the excited state to the ground state and vice versa. On the other hand, nonlinear optical processes with the χ⁽²⁾ terms can also be simulated successfully. The squeezed state of the NVE can be generated in this manner of the hybrid system. Comparing with the previous schemes with the nonlinear coupling of solid-state systems [4, 5], our scheme focuses on a new quantum system (involving NVE) with nonlinear coupling and our nonlinear strength can be enhanced using currently available technology. We will discuss all sources of decoherence and dissipation of the hybrid system in our following description.

2. The model

As schematically shown in Fig. 1, we consider a hybrid setup consisting of a NVE, a TLR and a SQ. The NVE contained in a single diamond crystal is placed at the antinodes of the magnetic field of the full-wave mode of the TLR with the geometric length L₀, distributed inductor L, and capacitance C. Similarly, the SQ interacts with the TLR by the magnetic flux through the loop enclosed in the SQ. The state of the SQ can be separately controlled by the gate voltage V_g through a gate capacitance C_g. The combined NVE-TLR-SQ system we study is governed by the Hamiltonian

H_{tot} = H_{NVE} + H_{FQ} + H_{TLR} + H_{STI} + H_{NTI},

in which H_NVE, H_FQ and H_TLR are the Hamiltonians of NVE, FQ and TLR, respectively. H_STI corresponds to the interaction between the SQ and TLR, and H_NTI represents the coupling between the NVE and the TLR.

Fig. 1 (a) Schematic of the hybrid system, where the NVE is placed in the antinode of the TLR coupled to the SQ through magnetic field induced by its quantized current. (b) The level structure of the NVE.

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The ground state of a NV is spin triplet with a zero-field splitting D_gs = 2.87 GHz between m_s = 0 and the nearly degenerate sublevels m_s = ± 1 in the absence of external magnetic field [39]. In our proposal, the logic states |g〉 and |e〉 of the NVE are denoted by the states |³A, m_s = 0〉, and |³A, m_s = −1〉, respectively. So the Hamiltonian of a single NVE reads (in units of ħ = 1),

H_{NVE} = δ S_{z},

where δ ≈ D_gs − γ_eB₀ is the energy gap between the ground state sublevels m_s = 0 and m_s = −1, with the external magnetic field B₀ and γ_e the the gyromagnetic ratio of an electron.

S_{z} = \frac{1}{2} \sum_{j = 1}^{N} τ_{j_{z}}

is the collective spin operator for the NVE with the Pauli spin operators τ_z = |e〉〈e| − |g〉〈g|, τ₊ = |e〉〈g| and τ₋ = |g〉〈e|. By using the definition of the collective operators of the NVE, we can easily find that these operators satisfy the SU(2) angular momentum commutation relations [S_z, S_±] = ±S_± and [S₊, S₋] = 2S_z, where the Casimir invariant is S² = N(N/2 + 1)/2.

In the present system, a superconducting quantum interference devices (SQUID) is employed, in which the SQ is coupled to a large superconducting reservoir through two identical Josephson junctions with capacitance C_J and Josephson coupling energy E_J. The SQ with zero and one extra Cooper pair in the superconducting island represented by the charge states |0〉 and |1〉, respectively, can act as an effective two-level artificial atom. The SQUID configuration of the SQ allows one to apply external flux Φ_e to control the total effective Josephson coupling energy. In the charge representation, the Hamiltonian of the SQ reads [40]

H_{SQ} = - \frac{1}{2} E_{C} (1 - 2 n_{g}) σ_{z} - E_{J} cos (π \frac{Φ_{e}}{Φ_{0}}) σ_{x},

where

E_{C} = \frac{2 e^{2}}{C_{Σ}}

is the single Cooper-pair charging energy of the SQ, and

n_{g} = \frac{C_{g} V_{g}}{(2 e)}

is the amount of gate charge induced by the gate voltage V_g with C_J Jopsephson junction capacitance and C_g gate capacitance. C_Σ = 2C_J + C_g is the total capacitance, Φ₀ = π/e is the flux quantum. σ_z = |1〉〈1| − |0〉〈0|, σ_x = |1〉〈0| + |0〉〈1| are the usual Pauli operators. The state of the SQ can be separately controlled by the gate voltage V_g through a gate capacitance C_g.

The coupling between the SQ and the TLR comes from the quantized external flux through the total effective area S of the SQUID. The SQUID is also integrated at the antinode of the magnetic field in the transmission-line resonator, where the electric field is zero or can be neglected. The quantized flux induced by the quantized current I_q of the TLR is

Φ_{q} = \frac{μ_{0} S I_{q}}{2 π d}

with d being the distance between the SQ and the TLR. At the antinode, the current in the TLR reaches its maximal amplitude

I_{q} = - i Σ_{k} I_{0}^{(k)} (a_{k}^{†} - a_{k})

, where a_k is the kth mode of the TLR,

I_{0}^{(k)} = \sqrt{\frac{ℏ ω_{k}}{L l}}

[19] with l being the length of the TLR and ω_k the frequency of the TLR. Considering that the SQ is placed at the point where the amplitude of the quantized current takes its maximal value, we can achieve the quantized flux as

Φ_{q} = - i \sum_{k} ϕ_{k} (a_{k}^{†} - a_{k}),

ϕ_{k} = \frac{μ_{0} S I_{0}^{(k)}}{2 π d} .

There is only one mode that couples to the SQ and the quantized flux felt by the SQUID becomes

Φ_{q} = - i ϕ_{b} (a^{†} - a) .

With the total biased flux, the effective Josephson coupling energy of the SQUID becomes

E_{J} (Φ_{q}) = - E_{J} cos (π \frac{Φ_{e} + Φ_{q}}{Φ_{0}}) .

Thus we could achieve the system combining the SQ and TLR by the Hamiltonian H₁ = H_SQ + H_TLR + H_STI as

H_{1} = ω_{a} a^{†} a - \frac{1}{2} E_{C} (1 - 2 n_{g}) σ_{z} - E_{j} cos (\frac{π Φ_{e} + Φ_{q}}{Φ_{0}}) σ_{x} .

It is obvious that the Josephson coupling energy in Eq. (9) is a nonlinear function of Φ_q. At some external biased points, for instance, Φ_e = mΦ₀, with m an arbitrary integer, the Josephson coupling energy in Eq. (9) becomes (−1)^m+1E_J cos(πΦ_q/Φ₀). If we use the eigenenergy basis of the SQ Hamiltonian to simplify the Hamiltonian in Eq. (9) and expand the Josephson coupling energy to the second order in Φ_q/Φ₀, Eq. (9) is simplified to (in units of ħ = 1),

H_{2} = \frac{1}{2} Ω {\tilde{σ}}_{z} + ω_{a} a^{†} a + η (a^{† 2} {\tilde{σ}}_{-} + a^{2} {\tilde{σ}}_{+}),

where the transition frequency of the SQ and the coupling coefficient between the TLR and the SQ in the nonlinear interaction regime are

Ω = \sqrt{E_{c}^{2} {(1 - 2 n_{g})}^{2} + 4 E_{J}^{2}},

η = {(- 1)}^{m + 1} \frac{E_{J} π^{2} ϕ_{b}^{2}}{2 Φ_{0}^{2}} cos θ,

respectively. a (a^†) is the annihilation (creation) operator of the full-wave mode of the TLR with the equilibrium frequency ω_a. We have used in Eq. (10) the new basis

| \tilde{0} 〉 = cos (\frac{θ}{2}) | 0 〉 + sin (\frac{θ}{2}) | 1 〉,

| \tilde{1} 〉 = - sin (\frac{θ}{2}) | 0 〉 + cos (\frac{θ}{2}) | 1 〉

with the mix angle θ = tan⁻¹[(−1)^m2E_J/(E_c(1 − 2n_g))]. Therefore, the corresponding Pauli operators are σ̃_z = |1̃〉〈1̃| − |0̃〉〈0̃|, σ̃₊ = |1̃〉〈0̃| and σ̃₋ = |0̃〉〈1̃|. Equation (10) includes an important nonlinear interaction which could be applied for an outstanding task in quantum mechanics and quantum optics, namely generating squeezed states [4, 5, 10, 21, 22]. Here we notice that this nonlinear coupling is very weak,which is reported as η = 2π × 0.25 kHz [21].

On the other hand, all the spins in NVE interact symmetrically with a single mode of the electromagnetic field because the mode wavelength is larger than the spatial dimension of the NVE if the spin ensemble is placed near the antinode of TLR’s field. The dynamics of the magnetic-dipole coupling between the TLR and the NVE is governed by the corresponding Hamiltonian

H_{NTI} = g (S_{+} a + S_{-} a^{†}),

which is a Jaynes-Cummings-type interaction with g the interaction strength between the NVE and the photon. Then the Hamiltonian of the whole system turns to be

H_{Tot} = \frac{1}{2} Ω {\tilde{σ}}_{z} + ω_{a} a^{†} a + δ S_{z} + η (a^{† 2} {\tilde{σ}}_{-} + a^{2} {\tilde{σ}}_{+}) + g (S_{+} a + S_{-} a^{†}) .

The last two terms of Eq. (16) shows two kinds of interactions, one of which is the quadratic coupling between the SQ and the TLR and the other is linear coupling between the NVE and the TLR. In following discussions, we mainly focus on the dispersive interaction with adiabatic elimination method by using a canonical transformation—the Fröhlich-Nakajima (FN) transformation [41]. In order to get to the effective nonlinear interaction between the NVE and SQ, the degrees of freedom of the TLR are eliminated. We will also show how this effective nonlinear interaction strength is enhanced by the factor of the total spin numbers in the NVE. As an application, we intend to use the nonlinear coupling to realize the squeezing of the NV, which is within reach of the currently available technology.

3. Enhanced strong nonlinear interaction

In terms of normalized collective operators ${\tilde{S}}_{+} = \frac{1}{\sqrt{N}} \sum_{j = 1}^{N} τ_{j_{z}}$ , we notice that the collective spin operator S̃₊ can create symmetric Dicke excitation states |n〉_P, and we encode the qubit of the NVE as

{| 0 〉}_{P} = | g_{1} g_{2} \dots g_{N} 〉

{| 1 〉}_{P} = {\tilde{S}}_{+} {| g 〉}_{N V} = \frac{1}{\sqrt{N}} \sum_{k = 1}^{N} τ_{+}^{k} | g_{1} g_{2} \dots g_{N} 〉 .

Hence, in this form we end up again with a two-level spin system, where strong coupling can be achieved for sufficiently large N. Thus we can rewritten the Eq. (16) as

H_{Tot} = \frac{1}{2} Ω {\tilde{σ}}_{z} + ω_{a} a^{†} a + δ {\tilde{S}}_{+} {\tilde{S}}_{-} + η (a^{† 2} {\tilde{σ}}_{-} + a^{2} {\tilde{σ}}_{+}) + \sqrt{N} g ({\tilde{S}}_{+} a + {\tilde{S}}_{-} a^{†}) .

In the dispersive regime, we consider the off-resonant case with the following conditions

Δ_{S} = | ω_{a} - δ | ≫ \sqrt{N} g,

Δ_{N} = | 2 ω_{a} - Ω | ≫ η,

which are also requisite by the usual second-order perturbation theory. The microwave photon is virtually excited under this situation. In terms of the FN transformation as

{\tilde{H}}_{Tot} = U_{R}^{†} H_{Tot} U_{R} = exp (- F) H_{Tot} exp (F)

, where F is the corresponding anti-Hermitian operator F for FN transformation adopts the following form

F = \frac{\sqrt{N} g}{Δ_{S}} (a^{†} {\tilde{S}}_{-} - a {\tilde{S}}_{+}) + \frac{η}{Δ_{N}} (a^{†^{2}} {\tilde{σ}}_{-} - a^{2} {\tilde{σ}}_{+}) .

Further transformation using the Baker-Campbell-Hausdorff formula

exp [- F] H_{Tot} exp [F] = H_{Tot} + [H_{Tot}, F] + \frac{1}{2!} [[H_{Tot}, F], F] + \dots

, expands the Hamiltonian H̃_Tot to the second order in

\sqrt{N} g / Δ_{S}

and η/Δ_N. Since the cavity mode only supports the virtual excitation of photons, we can neglect the terms with microwave photons and achieve an effective Hamiltonian

{\tilde{H}}_{Tot} ≃ \frac{1}{2} Ω {\tilde{σ}}_{z} + (ω_{0} + χ) {\tilde{S}}_{+} {\tilde{S}}_{-} + i λ_{eff} ({\tilde{S}}_{+}^{2} {\tilde{σ}}_{-} - {\tilde{S}}_{-}^{2} {\tilde{σ}}_{+}),

in which photon-induced dynamic energy shift of the SQ is neglected due to being a small quantity. Equation (23) shows clearly that, although no energy exchanges with the NVE, the microwave photon has still acted as a bus, generating long-range spin-spin interactions and the nonlinear NVE-SQ coupling. Here,

χ = \frac{N g^{2}}{Δ_{S}}

is the total photon-assisted long-range spin-spin interaction.

λ_{eff} = \frac{N g^{2} η}{Δ_{S}} (\frac{1}{Δ_{S}} + \frac{1}{Δ_{N}})

is the nonlinear NVE-SQ coupling constant. This nonlinear coupling λ_eff (S̃₊²σ̃₋ + S̃₋²σ̃₊) in the Hamiltonian H̃_Tot, to our knowledge, is new in quantum optics and condensed matter physics. For a single spin of NV, we know that

τ_{+}^{2} = 0

and

τ_{-}^{2} = 0

. Therefore, our achieved Hamiltonian only describes the coupling of the SQ to the collective excitation of the NVE.

When the SQ decays from the excited state to the ground state, two collective excitations of the NVE are achieved and vice versa. For large ensembles with few excitations our scheme is closely related to the bosonization procedure [42]. we find that for few excitations it is possible to identify S̃₊ = c^† and S̃₋ = c with bosonic creation and annihilation operators. By using the Holstein-Primakoff (HP) transformation [43] as

S_{+} ≃ \sqrt{N} c^{†}, S_{-} ≃ \sqrt{N} c, S_{z} ≃ c^{†} c - \frac{N}{2},

where the operator c^† (c) fulfills the approximate bosonic commutation relation [c, c^†] ≃ 1 in the low-excitation case, namely, in the limit of large N and low excitations, c^†c ≪ N. It means that the collective excitation governed by S_± behaves like a harmonic oscillator degree of freedom. Thus, the Hamiltonian H̃_Tot can be rewritten as

{\tilde{H}}_{Tot}^{'} ≃ \frac{1}{2} Ω {\tilde{σ}}_{z} + δ_{C} c^{†} c + - λ_{eff} (c^{2} {\tilde{σ}}_{+} - c^{†^{2}} {\tilde{σ}}_{-}),

where δ_C = ω₀ + Nχ.

From the effective Hamiltonian H̃′_Tot, one can find that this nonlinear coupling strength λ_eff is in direct proportion to the spin number N in ensemble. It implies that the coupling strength λ_eff can be enhanced greatly by increasing the value of N, as shown below. Physically, the effective Hamiltonian H̃′_Tot describes an interaction between the SQ and the pair of bosons excited from the NVE. The interaction represents that a pair of bosons is excited when the SQ decays from the excited state to the ground state and vice versa. This is also the reason why the coupling strength is in direct proportion to the spin number N in ensemble, which is the nonlinear effect different from traditional linear coupling proportion to $\sqrt{N}$ in ensemble [31–33]. This kind of nonlinear interaction might be useful for processing quantum information with continuous variables under strong coupling [6]. Moreover, the effective Hamiltonian H̃′_Tot also shows that the nonlinear optical processes with the χ⁽²⁾ term can be simulated successfully. Therefore, the quantum operations, such as squeezing, can be generated in our proposed hybrid systems. The details will be discussed in the next section.

To estimate the strength of the coupling between the NVE and the SQ, we consider currently available experimental parameters. Ω is allowed to be tunable in the range 2π × 10³ ∼ 10¹⁰ Hz [19] and the energy gap of the NVE could be also adjusted by external magnetic field. In the nonlinear interaction regime, the coupling strength is η = 2π × 0.25 kHz [21]. On the other hand, a single diamond crystal (3 × 3 × 0.5 mm³) containing the NVE is glued on top of TLR, with the distance to the silicon substrate less than 0.5 μm to ensure a maximal spin-resonator coupling. With the increase of the spin number N, the effective SQ-NVE coupling strength will be enhanced rapidly. If the coupled number of the NV centers in NVE is 10¹², the coupling between a single NV center and the TLR is 2π × 11 Hz and $\sqrt{N} g / Δ_{S} = η / Δ_{N} = 0.1$ . So we can estimate our nonlinear indirect interaction between the NVE and the SQ to be $λ_{eff} = \frac{N g^{2} η}{Δ_{S}} (\frac{1}{Δ_{S}} + \frac{1}{Δ_{N}}) ≃ \frac{{(\sqrt{N} g)}^{2}}{Δ_{S}} (\frac{η}{Δ_{N}}) = 0.01 \sqrt{N} g$ , and obtain λ_eff ≃ 2π × 0.1 MHz.

4. Squeezing

As discussed in the section III, the nonlinear optical processes with the χ⁽²⁾ term in Eq. (27) means that the essential quantum operations such as squeezing and parametric amplification can be implemented in our proposed hybrid SQ-TLR-NVE system. If we choose Ω = 2δ_C and shift to the rotating frame defined by $\frac{1}{2} Ω {\tilde{σ}}_{z} + δ_{C} c^{†} c$ , we obtain the following effective Hamiltonian

{\tilde{H}}_{R} ≃ i λ_{eff} (c^{2} {\tilde{σ}}_{+} - c^{†^{2}} {\tilde{σ}}_{-}) .

We now consider a spin-echo-like process in which we let the system evolve under the Hamiltonian in (28) for short periods of time Δt [5]. Between each of such time interval, we apply a quick π-pulse σ̃_x to flip the state of the charge qubit, so that the evolution for two periods can be simplified as e^−iH_sqt with H_sq = λ_eff (c² + c^†²)σ̃_x. If we initialize the charge qubit in the 〈σ̃_x〉 = 1 state, and repeat this procedure for M times, the evolution operator on the state of the NVE becomes

S (κ) = \frac{i κ}{2} (c^{2} - c^{†^{2}}),

with κ = Mλ_effΔt. This is a squeezing operator on the NVE with squeezing parameter κ. Assuming that the NVE initially in the state |0〉_P and using the transformation S^†(κ)cS(κ) = ccoshκ − c^† sinhκ, we obtain the variances in the two quadratures X₁ = (c + c^†)/2 and X₂ = (c − c^†)/2i,

Δ X_{1} = \sqrt{〈 X_{1}^{2} 〉 - {〈 X_{1} 〉}^{2}} = \frac{e^{- κ}}{2},

Δ X_{2} = \sqrt{〈 X_{2}^{2} 〉 - {〈 X_{2} 〉}^{2}} = \frac{e^{κ}}{2} .

As a result, the NVE evolves into the squeezed vacuum state

| κ 〉 = S (κ) | 0 〉 = \frac{κ}{2} (c^{2} + c^{†^{2}}) | 0 〉 .

It shows that the operators or the spins are squeezed fully for any spin number.

In order to estimate the practical squeezing efficiency, we also have to consider the influence from the noise in this hybrid SQ-TLR-NVE system. In the following, we will study the influence from the phase fluctuation and the amplitude fluctuation on the squeezing process, respectively. We firstly consider the phase fluctuation with Θ being the fluctuations in the phase of the pumping field with the effective Rabi frequency Ω_p.

From the Heisenberg equation [44], we can obtain the following stochastic multiplicative equation:

〈 \dot{Ψ} 〉 = (M_{0} - Θ M^{2}) 〈 Ψ 〉,

where

Ψ = [\begin{matrix} c^{+} c + c c^{+} \\ c^{2} \\ {(c^{+})}^{2} \end{matrix}], M_{0} = [\begin{array}{l} 0 & - 2 κ & - 2 κ \\ - κ & 0 & 0 \\ - κ & 0 & 0 \end{array}], M = [\begin{array}{l} 0 & 0 & 0 \\ 0 & 1 & 0 \\ 0 & 0 & - 1 \end{array}] .

With our specific initial condition 〈Ψ(0)〉 = {1, 0, 0}^T, the Laplace-transform solution has the following exact solution:

〈 c^{+} c + c c^{+} 〉 = \frac{- 2 κ}{2 π i} \int d z \frac{e^{z t}}{z (z + Θ) - 4 κ^{2}}

= e^{- Θ t / 2} (\frac{Θ sinh (ϰ t / 2)}{ϰ} + cosh (ϰ t / 2))

and

〈 c^{2} 〉 = \frac{Ω κ e^{- 2 Θ t}}{2 κ^{2} - Θ^{2} / 8} + \frac{(4 κ + 5 Θ) e^{(2 κ - \frac{3}{2} Θ) t}}{16 κ + 4 Θ} + \frac{(4 κ - 5 Θ) e^{(- 2 κ - \frac{3}{2} Θ) t}}{16 κ - 4 Θ},

where

ϰ = \sqrt{Θ^{2} + 16 κ^{2}}

. From the above two equations, we obtain the following formulas for the variance of the Hermitian amplitudes with laser phase fluctuations:

Δ X_{1} = \sqrt{\frac{e^{- Θ t / 2} [Θ sinh (ϰ t / 2) + 4 ϰ cosh (ϰ t / 2)]}{4 ϰ} + ξ}

Δ X_{2} = \sqrt{\frac{e^{- Θ t / 2} [Θ sinh (ϰ t / 2) + 4 ϰ cosh (ϰ t / 2)]}{4 ϰ} - ξ},

where

ξ = \frac{2 Θ κ e^{- 2 Θ t}}{8 κ^{2} - Θ^{2} / 2} + \frac{(4 κ - 5 Θ) e^{(- 2 κ - \frac{3}{2} Θ) t}}{32 κ - 8 Θ} - \frac{(4 κ + 5 Θ) e^{(2 κ - \frac{3}{2} Θ) t}}{32 κ + 8 Θ} .

In the limit Θ ≪ t⁻¹ ≪ Ω_p, where Ω_p = κ/t, the variances in the two quadratures ΔX₁ and ΔX₂ are given by

Δ X_{1} = \frac{\sqrt{(\frac{Θ t}{2}) e^{2 Ω_{p} t} + e^{- 2 Ω_{p} t}}}{2},

Δ X_{2} = \frac{e^{Ω_{p} t} \sqrt{1 - 2 Θ t}}{2} .

Similarly, the influence of the amplitude fluctuation can also be calculated as

Δ X_{1}^{'} = \frac{exp [2 Θ^{'} t - Ω_{p} t + 2 (e^{- Θ^{'} t} - 1)]}{2},

Δ X_{2}^{'} = \frac{exp [2 Θ^{'} t + Ω_{p} t + 2 (e^{- Θ^{'} t} - 1)]}{2},

where Θ′ is the fluctuations in the amplitude of the pumping field. We plot variances versus Ω_pt under different cases in the Fig. 2. It is shown that there are minimums of ΔX₁ and ΔX₂ as shown in Fig. 2(a), which means that to observe the squeezing in a noisy environment, it is important to accurately tune the values of Ω_pt. We also find that the phase fluctuation Θ affects the squeezed properties severely with increasing time t. On the other hand, in Fig. 2(b), the amplitude fluctuation Θ′ influences the squeezing efficiency nearly negligibly when Θ′/Ω_p ≤ 0.2.

Fig. 2 ((a) ΔX₁ versus Ω_pt under the different cases, where the soild red, dotted blue, dashed green, and dotdashed pink lines denote Θ/Ω_p = 0, Θ/Ω_p = 0.001, Θ/Ω_p = 0.01, and Θ/Ω_p = 0.1, respectively. The inset shows the ΔX₂ versus Ω_pt. (b) ΔX′₁ versus Ω_pt under the different cases, where the solid red, dotted blue, dashed green, and dotdashed pink lines denote Θ′/Ω_p = 0, Θ′/Ω_p = 0.1, Θ′/Ω_p = 0.2, and Θ′/Ω_p = 0.5, respectively. The inset shows the ΔX′₂ versus Ω_pt.

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Additionally, all sources of decoherence and dissipation of both the NVE and SQ should be considered. In current experiments, the electron spin relaxation time T₁ of the NV ranges from 6 ms at room temperature [45] to sec at low temperature. In addition, the dephasing time T₂ = 350 μs induced by the nearby nuclear-spin fluctuation has been reported [46]. But the latest experimental progress [47] with isotopically pure diamond sample has demonstrated a longer T₂ with T₂ = 2 ms. Current conservative experimental parameters of decoherence of the SQ show the relaxation rate and dephasing rate of the qubit as γ_q = 1 MHz and γ_ϕ = 10 MHz [16–18], respectively. On the other hand, the decoherence of the qubit brings insignificant affect on the squeezing efficiency due to the fact that the SQ colud always stay in the ground state with 〈σ_z〉 = 1 during the squeezing process [5, 22].

In the present hybrid system, the main obstacle on obtaining large squeezing efficiency comes from the decay rate of NVE [22]. Considering the typical decoherence, which is from the nuclear spins of ¹³C defects in the NVs, we can use isotopically purified ¹²C diamond through the purification technique [47]. The dipole-interaction between the redundant nitrogen spins and the NVs is another major decoherence. There are two different ways to reduce the influence. One is to improve the conversion rate which would reduce the linewidth of the NVs while maintaining the large collective coupling constants [31,38]. The other is to apply the external driving pulses to the electron spins on the redundant nitrogen atoms by using flip-flop processes, where the nitrogen spins are flipped by the spin-echo pulses on a time scale much faster than the flip-flop rate.

5. Conclusion

In summary, a hybrid system consisting of SQ, TLR, and NVE has been considered for achieving strong nonlinear interaction. Through the exchange of virtual microwave photons of the TLR, we have put forward a new nonlinear coupling between the NVE and the SQ. Our proposed hybrid system has described a nonlinear interaction between a pair of bosons and the SQ. In addition, we have also achieved nonlinear optical processes with the χ⁽²⁾ term, based on the collective excitations of the NVE. Therefore, the squeezed states can be generated.

Acknowledgments

This work is supported by the National Fundamental Research Program (No. 2013CB921803), the National Natural Science Foundation of China under Grants (No. 11104326, No. 11274351, and No. 11274351).

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Nonlinear coupling between a nitrogen-vacancy-center ensemble and a superconducting qubit

Abstract

1. Introduction

2. The model

3. Enhanced strong nonlinear interaction

4. Squeezing

5. Conclusion

Acknowledgments

References and links

Cited By

Figures (2)

Equations (44)

Optics Express