In the field of quantum information processing (QIP) and quantum manipulation (QM), the quantized electromagnetic field has always been considered the most reliable medium and means, such as in cavity quantum electrodynamics (QED) system and the superconducting circuits (SC) system (Xiang et al., 2013; Blais et al., 2020; Carusotto et al., 2020; Clerk et al., 2020; Haroche et al., 2020; Blais et al., 2021). However, with the rapid development of cryogenic and micro-nano processing technologies, traditional mechanical or acoustic devices have gradually re-entered the quantum world (O’Connell et al., 2010; Aspelmeyer et al., 2014; Forn-Díaz et al., 2019). In particular, quantum acoustic devices (QADs) have become among the most valuable quantum units owing to their unique advantages (Naber et al., 2006; Gustafsson et al., 2014; Schuetz et al., 2015). First, for the suppression of quantum noise, QADs can isolate environmental phononic noise at a much higher efficiency compared to other electromagnetic systems. Second, QADs are more convenient to design and fabricate because of their inherent larger dimensions (Schuetz et al., 2015; Kuzyk and Wang, 2018). Third, they originate from mature magnetostrictive or piezoelectric technologies, which can further improve their adaptive capacity to establish various hybrid quantum systems (Manenti et al., 2016; Knörzer et al., 2018). Finally, QADs can induce strong coupling to different qubits; this coupling mechanism is mainly similar to the ion (or atom) trap system and is naturally applicable to multiple QIP schemes, which were first proposed in ion trap systems (Andrew Golter et al., 2016a; Andrew Golter et al., 2016b). Therefore, an acoustic-based hybrid system can provide a promising platform to implement QIP and QM targets to ensure the efficient tailoring of spin–phonon and spin–spin interactions (Rabl et al., 2009; Bennett et al., 2013). Working as a quantum data bus or transducer, QADs can provide strong interactions in a wide range of qubits, especially at the single-quantum level (Li et al., 2020; Zhou et al., 2021; Leng et al., 2022; Zhou et al., 2022). Taking the surface acoustic wave (SAW) cavity, for example, strong coupling to different qubits can be achieved, with an estimated cooperativity C ∼ g ²/(γκ) of C ∼ 10–100 (Schuetz et al., 2015).

Recently, other kinds of artificial qubits, including nitrogen-vacancy (NV), silicon-vacancy (SiV), and hexagonal boron nitride (hBN) color centers have been introduced into the field of quantum science (Gao et al., 2015; Lemonde et al., 2018; Tan et al., 2022; Zhao et al., 2022). Without an additional trap, a SAW-based device may also induce strong stressful interactions directly to this type of solid-state spin, which increased interest for QIP and QM. This study proposes a general hybrid quantum system of an ensemble of collective spins coupled to a surface acoustic wave (SAW) cavity through a sideband design. Assisted by a dichromatic optical drive with a phase-dependent control, this spin ensemble effectively mimicked different types of long-range spin–spin interactions, namely the Lipkin–Meshkov–Glick (LMG) model with different types (Lipkin et al., 1965; Castaños et al., 2006; Morrison and Parkins, 2008; Ma and Wang, 2009; Zhang et al., 2017). Through a phase-dependent modification of this optical drive, this ensemble of spins undergoes quantum phase transitions (QPTs) because of the so-called spontaneous symmetry breaking (SSB). In addition, this phase-controlled scheme can also supply both the dynamical preparation of spin-squeezed state (SSS) (Kitagawa and Ueda, 1993; Wineland et al., 1994) and the adiabatic preparation of the multiparticle entangled ground states via QPTs (Vidal et al., 2004a; Vidal et al., 2004b); i.e., the Greenberger–Horne–Zeilinger (GHZ) or W-type states (Anders and Mølmer, 1999; Mølmer and Anders, 1999; Zheng, 2001; Unanyan and Fleischhauer, 2003). Given the goal of performing realistic QIP and QM, the results of our investigations are a fresh attempt using an acoustic control and may also provide general and useful applications.

Here we consider the basic proposal for this hybrid system and the coupling mechanism, which are illustrated in Figures 1A, B, C. Without a loss of generality, we take an ensemble of the NV centers as an example. In Figure 1A, an ensemble of solid-state spins is set near the surface of the SAW cavity with a fundamental frequency ω _m /2π ∼ 0.1–1.0 GHz. A phase-dependent dichromatic field (λ ∼ 700 nm) is applied to transmit the optional classical field (θ ₁, ω ₁, Ω₁) or (θ ₂, ω ₂, Ω₂). According to the previous investigation, this type of spin–phonon coupling is analogous to the ion trap system, and this SAW mode is similar to a harmonic oscillator potential as shown in Figure 1B. For an NV center, the energy-level design is plotted in Figure 1C, where |e(g)⟩ denotes the excited (ground) state with the energy split ω ₀ ≫ ω _m . We can obtain only the longitudinal coupling to each NV center via the SAW cavity, based on the expression H ̂ L ∼ ℏ g 0 ( a ̂ + a ̂ † ) | e 〉 〈 e | and coupling strength g ₀ (Andrew Golter et al., 2016b). Meanwhile, we apply the dichromatic optical drive to each NV spin obtain to the transition process |e⟩ ⇌|g⟩ near-resonantly. This sideband design is illustrated in Figure 1C. The identical red and blue detuning satisfies Δ ≡ ω ₂−ω ₀ ≡ ω ₀−ω ₁ ∼ ω _m .

Therefore, the systemic Hamiltonian for each NV spin (i.e., we consider the jth NV spin) is expressed as (ℏ = 1) H ̂ j = ω 0 σ ̂ z j / 2 + ω m a ̂ † a ̂ + g 0 a ̂ + a ̂ † σ ̂ z j + σ ̂ j + Ω 1 e − i ω 1 t + Ω 2 e − i ω 2 t + H.c. , (1) where σ ̂ z j ≡ | e 〉 j ⟨ e | − | g ⟩ j 〈 g | , σ ̂ j + ≡ | e 〉 j 〈 g | . For simplicity, we assume here that the collective spins are homogenous, and write the Hamiltonian of this whole system as H ̂ = ω 0 S ̂ z + ω m a ̂ † a ̂ + g a ̂ + a ̂ † S ̂ z + S ̂ + Ω 1 e − i ω 1 t + Ω 2 e − i ω 2 t + H.c. (2) in which the collective spin operators are defined as S ̂ x , y , z = ∑ j = 1 N σ ̂ x , y , z j / 2 and S ̂ ± = ∑ j = 1 N σ ̂ j ± , with the basic commutation relation: [ S ̂ i , S ̂ j ] = i ε ijk S ̂ k , [ S ̂ + , S ̂ − ] = 2 S ̂ z , and [ S ̂ z , S ̂ ± ] = ± S ̂ ± . We apply the unitary Schrieffer–Wolff transformation U ̂ SW = e − i p ̂ to Eq. 2, with the Hermitian operator p ̂ ≡ i η ( a ̂ † − a ̂ ) S ̂ z and the Lamb–Dicke-like parameter η = g/ω _m ∼ 0.1–0.2 (Rabl et al., 2010; Zhou et al., 2018; Li et al., 2020). We assume that the SAW is cooled sufficiently at extremely low ambient temperature so that this hybrid system satisfies the so-called Lamb–Dicke limit ( n ̄ + 1 ) η 2 ≪ 1 , where n ̄ = 1 / ( e ℏ ω m / k B T − 1 ) is the average number of the phonon for this acoustic mode for the environmental temperature T. For example, we assume that ω _m /2π ∼ 1.0 GHz; once T ≈ 1.0 K, we get n ̄ ≈ 20.3 ; if T ≈ 0.1 K, we also get n ̄ ≈ 1.6 . Therefore, the thermal phonon number can be compressed effectively as long as this hybrid system is at a low temperature. Applying the approximate relation e ± η ( a ̂ † − a ̂ ) ≃ 1 ± η ( a ̂ † − a ̂ ) to Eq. 2 we acquire the Hamiltonian in the interaction picture. As η ≪ 1, we can effectively rewrite the Hamiltonian in the interaction picture (IP) H ̂ IP ≃ S ̂ + e i ω 0 t × 1 + η a ̂ † e i ω m t − a ̂ e − i ω m t × Ω 1 e − i ω 1 t + Ω 2 e − i ω 2 t + H.c. (3) In this scheme, we assume the following relations: ν ≫ λ, |Δ|≫|Ω_1,2|, and {ν, |Δ|, |Δ ± ν|}≫ η|Ω_1,2| and can then eliminate this phonon mode adiabatically (according to Appendix A) (James, 1998; James and Jerke, 2012). Ignoring the items for the energy shift caused by this acoustic mode ( ∼ a ̂ † a ̂ S ̂ z ), we can get the general LMG model with the long-range spin–spin interactions. H ̂ Total eff ≈ A S ̂ z + B S ̂ x 2 + C S ̂ y 2 + D S ̂ x S ̂ y + S ̂ y S ̂ x , (4) where the relevant coefficients are A = 2 Δ − 2 η 2 Δ δ 2 δ 3 r 1 2 − r 2 2 , B = 2 ω m η 2 δ 2 δ 3 2 r 1 r 2 ⁡ cos Θ − r 1 2 − r 2 2 , C = 2 ω m η 2 δ 2 δ 3 − 2 r 1 r 2 ⁡ cos Θ − r 1 2 − r 2 2 , D = − 4 ω m η 2 δ 2 δ 3 r 1 r 2 ⁡ sin Θ .

The Lipkin–Meshkov–Glick (LMG) model was first proposed in the area of nuclear physics to describe monopole-monopole interactions and is one kind of solvable long-range spin–spin model (Lipkin et al., 1965). This model, not only obeys the conservation of angular momentum but also satisfies the Z 2 symmetry. Thus, many theoretical proposals have been described for simulating this spin model to explore new topics in physics. Different physical systems such as the ion-trap system (Zheng, 2001; Unanyan and Fleischhauer, 2003), the cavity QED system (Morrison and Parkins, 2008; Zhang et al., 2017), and the superconducting system (Tsomokos et al., 2008) have presented theoretical schemes and reported interesting results. The general LMG model may be expressed as H ̂ = ϵ S ̂ z + V S ̂ x 2 − S ̂ y 2 + W S ̂ x 2 + S ̂ y 2 , (5) and the relevant tunable parameters are ϵ, V, and W, respectively (Lipkin et al., 1965). Here, Eq. 5 satisfies the conservation of angular momentum, namely, [ H ̂ , S 2 ] with S 2 ≡ S ̂ x 2 + S ̂ y 2 + S ̂ z 2 = S ( S + 1 ) . In addition, we can achieve several types of LMG models by modifying the aforementioned parameters. For example, once V = 0 and W ≠ 0, we can first obtain isotropy-type interactions with H ̂ I = ϵ S ̂ z + W ( S 2 − S ̂ z 2 ) . This type of LMG Hamiltonian can be solved exactly in the representation of S ̂ z ; i.e., S ̂ z | m z 〉 = m z | m z 〉 , −S ≤ m _z ≤ S and |m _z ⟩ denote the eigenstate of S ̂ z . We can, therefore, obtain the relation H ̂ I | m z 〉 = E m z | m z 〉 and E m z = ϵ m z + W [ S ( S + 1 ) − m z 2 ] . This model mainly indicates the ferromagnetic order, with a ground state of |m _z = S⟩ or |m _z = −S⟩. When ϵ = 0 and V = W ≠ 0 (or V = −W ≠ 0), we obtain the one-axis twisting LMG model; i.e., H ̂ x , y = C x , y S ̂ x , y 2 with C _x = W + V and C _y = W−V. Here, H ̂ x , y belongs to the ferromagnetic order when C _x,y < 0; conversely, when C _x,y > 0, it obeys the antiferromagnetic order. Furthermore, when V ≠ W ≠ 0, we obtain the two-axis twisting LMG model; i.e., H ̂ T = ϵ S ̂ z + C x S ̂ x 2 + C y S ̂ y 2 .

First, according to the previous investigation of this long-range spin model, the isotropy LMG model is (see Appendix B) H ̂ Isotropy 1,2 = A 0 1,2 S ̂ z − B 0 1,2 S ̂ z 2 + B 0 1,2 S 2 . (6) In this Hamiltonian, S 2 ≡ S ̂ x 2 + S ̂ y 2 + S ̂ z 2 = S ( S + 1 ) is a conserved quantity, S = N/2 represents the maximum total angular momentum, and H ̂ Isotropy 1,2 corresponds to a phase-independent model. This also belongs to the ferromagnetic- (FM-) order Hamiltonian, with the unique and non-degenerate ground state namely, |Ψ⟩ _u = |↑↑⋯⟩ ≡|m _z = S⟩ or |Ψ⟩ _d = |↓↓⋯⟩ ≡|m _z = −S⟩, with spin number N (Zheng, 2001). Here, the coefficient A 0 1,2 breaks the original symmetry and eliminates the degeneracy of item − B 0 1,2 S ̂ z 2 . The unique ground state |Ψ⟩ _u or |Ψ⟩ _d is decided mainly by the parameter A 0 1,2 . For example, if we can obtain H ̂ Isotropy 2 and A 0 2 > 0 , the ground state is |Ψ⟩ _u ; however, for H ̂ Isotropy 1 and A 0 1 > 0 , the ground state is |Ψ⟩ _d .

Second, the more interesting and novel point lies in the phase-dependent control of this proposal, which we discuss briefly. Here, we assume Ω 1,2 = r 1,2 e i θ 1,2 and Θ = θ ₁+θ ₂. To study the effects caused by the phase of this dichromatic optical drive, we assume r ₁ = r ₂ = r for simplicity, to get H ̂ General LMG ≈ − G T 1 − cos Θ S ̂ x 2 + 1 + cos Θ S ̂ y 2 + sin Θ S ̂ x S ̂ y + S ̂ y S ̂ x , (7) with G T = 4 ω m r 2 η 2 δ 2 δ 3 .

For single solid-state spin, we first define | ± 〉 x = ( | ↑ 〉 ± | ↓ 〉 ) / 2 and | ± 〉 y = ( | ↑ 〉 ± i | ↓ 〉 ) / 2 . We can then make the brief list shown in Table 1. For example, for the first case, namely, case (1), with Θ = 0, we obtain the one-axis twisting LMG model along the y direction, H ̂ y LMG = − G y S ̂ y 2 . (8) When G y = 8 ω m r 2 η 2 δ 2 δ 3 > 0 , the Hamiltonian H ̂ y LMG presents the ferromagnetic (FM) order, with a ground state corresponding to |m _y = ±N/2⟩ = |±±⋯⟩ _y ≡|±⟩ _y |±⟩ _y ⋯|±⟩ _y ; however, when G _y < 0, the Hamiltonian H ̂ y LMG is the antiferromagnetic (AFM) order, with a ground state of |m _y = 0⟩ (N is an even number) or |m _y = ±1/2⟩ (N is an odd number). In addition, according to our previous investigation on this type of interaction, we can obtain an adiabatic transition process from the initial disentangled ground state (governed by H ̂ Isotropy 1,2 ) to the N-particle Greenberger–Horne–Zeilinger (GHZ)-type entangled state (Zhou et al., 2018).

While for case (2), when Θ = π, we can also get another one-axis twisting LMG model along the x direction, H ̂ x LMG = − G x S ̂ x 2 . (9) Equivalently, when G x = 8 ω m r 2 η 2 δ 2 δ 3 > 0 , the Hamiltonian H ̂ x LMG also shows the ferromagnetic (FM) order, with a ground state of |m _x = ±N/2⟩ = |±±⋯⟩ _x ≡|±⟩ _x |±⟩ _x ⋯|±⟩ _x ; however, when G _x < 0, the Hamiltonian H ̂ x LMG is of the antiferromagnetic (AFM) order, with a ground state of |m _x = 0⟩ (N is an even number) or |m _x = ±1/2⟩ (N is an odd number).

Then, for case (3), as illustrated in the third and fourth lines of Table 1, when Θ = ±π/2, we obtain a novel transverse-type LMG model (in the x−y plane) as follows: H ̂ T ± = − G T S ̂ x ± S ̂ y 2 . (10) This type of spin–spin interaction belongs to the one-axis twisting model. For example, if we define a linear transformation in this model; i.e., S ̂ + = 2 S ̂ + e i ϑ , S ̂ − = 2 S ̂ − e − i ϑ , and S ̂ x = ( S ̂ + + S ̂ − ) / 2 , we get H ̂ T ± = − G T S ̂ x 2 , with the “±” signs ϑ = ∓π/4, respectively. In this new x representation, we can also obtain an equivalent one-axis twisting LMG model along the x direction. The physical mechanism of this model is also similar to that in case (2), and its FM and AFM phases are governed by the sign of G _T ; i.e., G _T > 0, the Hamiltonian H ̂ T ± denotes the FM order, with a ground state of |m _x = ±N/2⟩ = |±±⋯⟩ _x ; but while G _T < 0, H ̂ x ± stands for the AFM order, with a ground state of |m _x = 0⟩ (N is an even number) or |m _x = ±1/2⟩ (N is an odd number).

Thus, despite the different representations, we can always get the one-axis twisting LMG model, which can also lead to the preparation of the spin-squeezed state (SSS) through a dynamic process. Two main definitions have been proposed to describe the squeezed degree of this spin ensemble, (Sidles et al., 1995). First, Masahiro Kitagawa and Masahito Ueda defined ξ S 2 = 4 min ( Δ J ̂ n ⃗ ⊥ 2 ) N in 1993. Another similar definition, the metrological spin-squeezing parameter ξ R 2 = N ( Δ J ̂ ⊥ ) 2 ⟨ J ̂ S ⟩ 2 , was introduced by Wineland et al. in 1992 and 1994 (Wineland et al., 1992; Wineland et al., 1994). According to our previous investigation, both definitions are valid and reliable for describing the SSS; moreover, we also reported that ξ S 2 ≤ ξ R 2 (Li et al., 2020). For simplicity, we plot the dynamical ξ R 2 (Kitagawa and Ueda, 1993) in Figure 2, by solving the following master equation numerically. Under the realistic condition considering both decay and dephasing factors (Γ_dc and Γ_dp), the corresponding dynamical process of this whole system is dominated by the master equation: d ϱ ̂ d t = i ϱ ̂ , H ̂ General LMG + Γ dc 2 S ̂ − ϱ ̂ S ̂ + − S ̂ + S ̂ − ϱ ̂ − ϱ ̂ S ̂ + S ̂ − + Γ dp 2 S ̂ z ϱ ̂ S ̂ z − S ̂ z S ̂ z ϱ ̂ − ϱ ̂ S ̂ z S ̂ z . (11)

From which we note that this general LMG model can always engineer collective spins into the SSS dynamically at time ∼ 0.03 G T (N = 100), or ∼ 0.05 G T (N = 50), or ∼ 0.08 G T (N = 20). In addition, governed by this general LMG model, no matter the special phase Θ or general Θ, the collective spins will be equivalently and efficiently engineered into the SSS.

Utilizing the standard average-field approximation, we can determine the expected value of the collective spin components via the definition ⟨ S ̂ i ⟩ = Tr ( ρ S ̂ i ) and then write a group of time-dependent differential equations for the total angular momentum through the basic relation d ⟨ S ̂ i ⟩ / d t = Tr ( ρ ̇ S ̂ i ) . We introduce an equivalent normalization transformation: ⟨ S ̂ x ⟩ = N X , ⟨ S ̂ y ⟩ = N Y , and ⟨ S ̂ z ⟩ = N Z to this proposal. To achieve the theoretical results, we also rewrite Eq. 11 for simplicity. d ρ ̂ d t = i ρ ̂ , H ̂ Θ + γ dc 2 S ̂ − ρ S ̂ + − S ̂ + S ̂ − ρ − ρ S ̂ + S ̂ − + γ dp 2 S ̂ z ρ S ̂ z − S ̂ z S ̂ z ρ − ρ S ̂ z S ̂ z , (12) in which we assume that all spins are homogenous and that their relative dephasing and decay rates are uniform factors; i.e., we assume that γ _dc ∼Γ_dc/(−G _T ) and γ _dp ∼Γ_dp/(−G _T ). By discarding the quantum fluctuations of S ̂ x , y , z , we effectively obtain a group of semiclassical equations: X ̇ = 2 cos Θ + 1 Y Z + 2 ⁡ sin Θ − γ dc X Z − γ dp X / 2 , Y ̇ = 2 cos Θ − 1 X Z + − 2 ⁡ sin Θ − γ dc Y Z − γ dp Y / 2 , Z ̇ = − 4 ⁡ cos Θ X Y + − 2 ⁡ sin Θ + γ dc X 2 + 2 ⁡ sin Θ + γ dc Y 2 . (13)

Together with the total spins’ conservation relation X ²+Y ²+Z ² = 1, we obtain the analytical solutions of the relevant order parameters, i.e., X, Y, and Z, by solving this group of equations. Then, we can determine the analytical solutions as (1) the trivial solution X = Y = 0, and Z = ±1 for the normal phase; and (2) the non-trivial solutions Z = − γ dp 2 γ dc , X = ± 1 − Z 2 cos Θ + 1 2 , Y = − X sin Θ cos Θ + 1 . (14) Assuming γ _dp ≈ 0.1γ _dc, we get Z = −0.05 and plot the numerical average-value order parameters X and Y in Figure 3. Owing to the Z ₂ symmetry of this general LMG model H ̂ Θ , we can obtain the positive and negative values symmetrically with X _± and Y _±. From which, we first note several critical points among the varying region [−2π, 2π].

More specifically, when Θ approaches points such as 0, ± 2π, and even 2kπ (k is any integer number), X ₊ → 1, X ₋ → −1, and Y _± → 0 simultaneously. This type of QPT is mainly induced by the transformation of the representation; i.e., z ↔ x or x ↔ y. We consider this to correspond to phase-induced spontaneous symmetry breaking (SSB). When Θ → ±π, we get a different QPT effect, namely, X _± → 0; however, its n-order derivative at this point ( ∂ n X ± / ∂ Θ n ) ± π is discontinuous with n ≥ 1. More interestingly, around these points, Y ₊ and Y ₋ will suddenly turn over to each other and both Y _± and ( ∂ n Y ± / ∂ Θ n ) ± π are discontinuous at this type of critical point. Finally, regarding other critical points, when Θ is modified near points such as ± π/2, ± 3π/2, ⋯, a synchronous crossing phenomenon will occur; i.e., X ₊ = Y ₊ and X ₋ = Y ₋. This result corresponds to the mixture-type LMG model (10), and it is also basically induced by the so-called SSB.

Therefore, no matter in case (1) and in cases (2) and (3), these long-range spin–spin interactions belong to the one-axis twisting LMG model under their different y, x, and x representations. This type of spin model can also play an important role in the preparation of multiparticle entanglement. Starting from the original and general LMG model (4), we briefly discuss this topic.

Initially, we can easily get the isotropy LMG model such as Eq. 6, which belongs to the FM order. Its ground state is the disentangled and unique state |m _z = S⟩ or |m _z = −S⟩, which is also decided by the sign of A 0 1,2 . The relevant results are plotted in Figure 4. Next, when we modify the amplitude of this classical field adiabatically with r _1,2 → r and select the phase θ _1,2 with Θ = θ ₁+θ ₂, we obtain not only the SSB-induced QPTs but also efficient adiabatic passage for engineering collective spins into the entangled ground state; i.e., the GHZ or W states. As illustrated in Figure 4, if this GS transition corresponds to the FM↦FM, we can get the entanglement passage for the GHZ state. However, when the GS transition is FM↦AFM, we get another entanglement strategy for the W-type GS. Because these types of investigations have been performed previously, these results confirm the feasibility of this scheme for the preparation of entanglement (Zhou et al., 2018).

In this proposal, we first considered an ensemble of solid-state spins located on the surface of the SAW cavity with a fundamental frequency of ω _m /2π ∼ 0.1–1.0 GHz (Schuetz et al., 2015). SAW-based quantum devices can induce strong or ultra-strong interactions in many kinds of atoms, spins, and even other artificial atoms, as shown in Table 2. Taking the nitrogen-vacancy (NV) spins for example, we can define the ground state |0⟩ ≡|g⟩ and excited state |E _y ⟩ ≡|e⟩, with energy-splitting ω ₀/2π ∼ 470 THz. The estimated coherent spin–phonon coupling at the single-quantum level is approximately g ₀/2π ∼ 100 kHz; thus, we can determine the collective coupling strength g ≈ N g 0 (Andrew Golter et al., 2016a; Andrew Golter et al., 2016b). The coherent spin–phonon coupling at the single-quantum level of a single quantum dot (QD) can reach g ₀/2π ∼ 100 MHz. For both different solid-state spins, their cooperativity in a SAW-based hybrid system is C ∼ g ²/(γκ) ∼ 10–100, which also belongs to the strong-coupling region (Schuetz et al., 2015).

Utilizing a hybrid system of collective spins coupled to a SAW cavity, we study a proposal for simulating the general long-range LMG model with a phase-dependent control. Our analysis and discussion show that this scheme can not only ensure the generation of the SSS via a dynamical evolution governed by the general one-axis twisting interactions but also supply a potential route toward multiparticle GHZ or W-type states through their FM or AFM phase transitions. We also studied the open-system critical behavior of this general LMG model by using the average-field method. For the target of carrying out realistic QIP and QM, our findings may be considered a fresh attempt by using an acoustic control, especially a phase-dependent control. Moreover, this proposed system might have various applications.